Information Preservation in Quantum Gravity

Lawrence B. Crowell · Mar 26, 2008

This is a second part to the post on the Fubini-Study metric and the Hawking effect. The Fubini-Study metric defines the set of projective rays in the complex space C^{n+1}, or CP^n~\subset~C^{n+1}. The Fubini-Study metric is then for n = 1 the Bloch sphere and further defines quantum entanglements and Berry phases.

A Hermitian function or differential form in C^{n+1} defines a unitary subgroup U(n+1)~\subset~ GL(n+1,C). The Fubini-Study metric is invariant under scaling under group actions of such a U(n+1). Hence the space, a Kahler manifold, is homogeneous. Hence any two Fubini-Study metrics are isometric under a projective automorphism of CP^n.

The definition of a projective space is the set of elements z_i~=~Z_i/Z_0, where the magnitude of Z_0 is artitrary, or more formally

 CP^n~=~\big{\{z_i,~i=1\dots n}:~z_i~=~Z_i/(Z_0~\ne~0)\big} 

A point or in CP^n is a line or ray in C^{n+1}. An arbitrary vector in the projective space may then be represented in the "bra-ket" notation of quantum mechanics as

 |\psi\rangle~=~\sum_{i=1}^\infty z_i|e_i\rangle~=~[z_1:~z_2:~\dots~z_n] 

where the square braket notation is commonly used for projective spaces. The basis vectors |e_i\rangle are the basis vectors for the Hilbert space {\cal H}~=~C^{n+1}. There there are two such vectors |\psi\rangle,~|\phi\rangle, then the quanatum mechanical overlap defines the distance between the two points in CP^n, which in the Hilbert space is a plane. The modulus square of the quantum overlap appropriately normalized

 \frac{\langle\phi|\psi\rangle\langle\psi|\phi\rangle}{\langle\psi|\psi\rangle\langle\phi|\phi\rangle}~=~cos(\theta(\psi,~\phi)) 

defines the distance between the two as \theta(\psi,~\phi). For n = 1 this is the angle on the Bloch sphere, and defines a general quantum angle.

The overlap between a vector |\psi\rangleand its infinitesimal displacement |\psi~+~\delta\psi\rangledefines the Fubini-Study metric on C^n with the proper normalization

 ds^2~=~\frac{\langle\delta\psi|\delta\psi\rangle}{\langle\psi|\psi\rangle}~=~\frac{\langle\delta\psi|\psi\rangle\langle\psi|\delta\psi\rangle}{|\langle\psi|\psi\rangle|^2} 

With the coordinate notation for projective spaces the Fubini Sudy metric is

 ds^2~=~\frac{1}{2}\frac{[z_i,~dz_j][{\bar z}^i,~d{\bar z}^j]}{(z_k{\bar z}^k)} 

where [z_i,~w_j] is a commutator. We might think of the basis elements of the Hilbert space as a Fock basis and the components z_i as due to the application of raising and lowering operators or a field amplitude of the form A_i~=~a^\dagger_ie^{i\theta}~+~ia_ie^{-i\theta}. The projective space is a form of algebraic variety, and these commutators are defined as Grassmannian varieties. The form of this metric indicates that the space is a Kahler manifold where the Ricci curvature is given by a potential \Phi

 R_{ij}~=~\frac{\partial^2\Phi}{\partial z_i\partial{\bar z}_j} 

which is proportional to the metric.

For any two dimensional subspace of CP^1~\subset~CP^n, the space is the block sphere of two real dimensional where the Fubini-Study metric reduces to

 ds^2~=~\frac{dzd{\bar z}}{(1~+~z^2)(1~+~{\bar z}^2)}~=~d\theta^2~+~sin^2\theta d\phi^2 

with the state |\psi\rangle~=~cos(\theta)|0\rangle~+~e^{i\phi}\sin(\theta)|1\rangle, which is the single qubit state. This may then be generalized to higher dimensional kahler manifolds for hexacode and higher quantum codes and algebras.

Lawrence B. Crowell

Lawrence B. Crowell · Mar 26, 2008

The last equation I quoted in this lastest thread should read

 T~=~\frac{\hbar c^3}{8\pi GM k}. 

That is what you get when you do math on an editor instead of on paper.

L. Crowell

Lawrence B. Crowell · Mar 27, 2008

A slice through spacetime will contain some set of black holes and likely have some cosmological event horizon. Each of these event horizons will then satisfy a quadratic polynomial for the coordinates specified in that region. Since the event horizon is an invariant it does not matter what coordinates are used, so long as the spacetime has some consistent set specified by some coordinate condition or the spacetime analogue of a gauge. Hence there now exists a subspace of the universe {\cal P}, a set of all projective rays that join at {\cal I}^\pm. Event horizons are those which satisfy a certain polynomial condition P(x_i)~=~0 which identifies them as event horizons. This is then a projective variety defined by a set of homogenous polynomials \cal S, such that some points of \cal P satisfy

 Z({\cal S})~=~\{x~\in~{\cal P}|P(x)~=~0,~\forall~P~\in~{\cal S}\} 

This subset of the projective space is a projective algebraic set,where the irreducible set defines a projective algebraic variety. This set has the Zariski topology by declaring all algebraic sets to be closed.

A Zariski topology is a nonHausdoff topology for algebraic varieties, which describes their algebraic content and with weak geometric content. Zariski topology does occur in general relativity. Gravitation has the group structure SO(3,~1)~=~Z_2\times SL(2,~C). The group SL(2,~C) may in turn be written as

 SL(2,~C)~=~SU(1,~1)\times SU(2). 

The algebra for this is then su(2) given by the standard Pauli matrices \sigma_{\pm},~\sigma_3 and su(1,~1) has the elements \sigma_{\pm},~\tau_3~=~i\sigma_3. The latter change gives the pseudoEuclidean nature to spacetime.

Now consider a connection one-form

 A~=~A^+\sigma_+~+~A^3\sigma_3, 

and a gauge transformation determined by the group action of g~\in~{\cal G}, g~=~e^{i\lambda\sigma_3}. The gauge transformed connection is then

 A^\prime~=~g^{-1}Ag~+~g^{-1}dg~=~e^{-2\lambda}A^+\sigma_+~+~A^3\sigma_3 

where d\lambda~=~A^3. Thus \lambda is a parameterization of the gauge orbit for this connection A^'~=~A(\lambda). This leads to the observation

 \lim_{\lambda\rightarrow\infty}A(\lambda)~\rightarrow~A^3\sigma_3, 

where A^+\sigma_+~+~A^3\sigma_3 and A^3\sigma_3 have distinct holonomy groups and thus represent distinct points in the moduli space \cal M. However by the last rquation we must have

 F_\mu(A^+\sigma_+~+~A^3\sigma_3)~=~F_\mu(A^3\sigma_3), 

and similarly for any gauge invariant function. Hence there exist two distinct point in the moduli space that define the same set of gauge invariant functions. Hence there does not exist a measure over these two points that separates them, and \cal M is then nonHausdorff and has Zariski topology.

This moduli space, a space of gauge equivalent connections, indicates the underlying topology for curvatures. In a b-completeness sense the group action could be g~=~e^{i\lambda\sigma_3~+~\alpha}, for \alpha a constant. The gauge connections for \lambda~\rightarrow~\infty converge to e^{-2\alpha}A^+\sigma_+~+~A^3\sigma_3, which for a given \alpha can describe a set or congruence of null geodesics. A Cauchy-like sequence of \lambda_n can describe a set of geodesic which converge to the null set. In this manner the underlying topological structure of gravitation has this nonHausdorff topology.

This structure then indicates that quantum gravity if more fundamentally described by algebraic varieties, with minimal geometric content. An abstract algebraic variety is a generalization of a scheme, which geometrically is a correspondence with some class of rings. In this case the polynomials define some ring. A scheme is a space with a local ring structure such that every point has a neighborhood, a locally ringed space, isomorphic to a ring spectrum. An algebraic variety is a scheme whose structure sheaf is a sheaf of K-algebras with the property that the rings R that occur above domains which are all finitely generated K-algebras, i.e., quotients of polynomial algebras by prime ideals.

The system works over any field F, and the topology permits one to piece together varieties on open sets and their definition permits them to be on a projective space. This leads to the topic of sheaf cohomology in algebraic geometry. Another mathematical issue of particular interest here is that a special case of an algebraic variety is an algebraic curve. This leads to the Goppa code. For X non-singular, or there exist points x_1,~x_2,~\dots,~x_n fixed among the points of X defined over F, and that G is a divisor on X, also defined over F. There is a finite-dimensional subspace L(G) of the function field of X, consisting of the rational functions f on X with zeroes and poles subject to G. This means that G, which is a formal sum of points of X over the algebraic closure of F, bounds the divisor made up of the zeroes and poles of f, counted with appropriate multiplicity.

An example of an algebraic variety is an algebraic curve of unity dimension. Simple examples are ellipses, conic sections and the more enigmatic elliptical curves. A curve has at most a finite number of singular points, where if that is zero the curve is nonsingular. This requires that the curve exists in a projective space and be over an algebraically closed field. The theory of nonsingular curves over complex numbers is equivalent to the theory of Riemannian surfaces, where the genus of the curve is that of the two-manifold. This genus is then a result of the Riemann-Roch theorem.

This defines structures which are codes, which preserve the number of quantum bits. To preserve this information it must process this information according to an error correction code, similar to those used in computer data transmission.

What has been developed here is commesurate with a Goppa code, which is a linear code constructed from algebraic curves over a finite field F. Let the curve C be nonsingular with a set of fixed points p_i, i~=~1.~\dots~,n. There then exists a divisor F on the curve C over the field F. Given rational functions f over the curve with poles and zeros on G, this defines a space L(G) of finite dimension defined as the function field of C. Thus the divisor G is a formal sum of point on C that bounds a divisor made from the poles and zeros of f, and is algebraically closed in F. Thus L(G) has some basis of functions f_1,~\dots,~f_k over the field, where the Goppa codes is found by the evaluation of the fixed points p_i, eg. f_i(p_1),~f_i(p_2),~\dots,~f_i(p_n), for i~=~1,~\dots,~k. This defines a vector space f^n spanned over F. This is then given by the map

 \mu:L(G)~\rightarrow~F^n, 

by f~\mapsto~f(p_1),~\dots,~f(p_n). For a divisor defined as D~=~\sum_{i=0}^np_i the Goppa codes is then represented as c(D,~G)~=~L(G)/ker(\mu).

For a function f~\in~ker(\mu), for f(p_i)~=~0 the function as a divisor is less than D and so ker(\mu)~=~L(G~-~D). For f~\in~L(G~-~D) since p_i~<~G then div(f)~>~D it is clear that f(p_i)~=~0. A distance between code words is the Hamming weight of \mu(f) called d. Thus f(p_i)~=~0 for n~-~d_i p_is. The points are designated p_{i_1},~\dots,~p_{i_{n-d}}. Thus

 f~\in~L(G~-~p_{i_1}~-~\dots~-~p_{i_{n-d}}), 

and that div(f)~<~ G~-~p_{i_1}~-~\dots~-~p_{i_{n-d}}. By taking the degree of these on both sides the degree of G then satisfies d~\ge~n~-~deg(G) The dimensions of the Goppa code is then k~=~L(G)~-~L(G~-~D). This result is the Riemann-Roch theorem.

The Riemann-Roch theorem indicates how many moduli exist. A moduli space is a space of solutions that defines gauge equivalent gauge vectors. The moduli space is the set of all self-dual connections (instantons) with degrees of freedom removed by gauge conditions. This space is \omega_{SD}/{\cal G}~=~{\cal M}_{mod}. The number of metric moduli and conformal Killing vectors of the moduli space is obtained with the Riemann-Roch theorem. Given an operator {\cal O} that is covariantly constant on a geodesic,

 \partial g_{ij}~=~ (2\delta\omega~-~ \nabla_k\delta \sigma))g_{ij}~-~ 2({\cal O}\delta \sigma)_{ij}, 

where \omega is the conformal factor, \sigma is a spatial parameter. Here the metric is defined on the complexified coordinates z,~{\bar z}. {\cal O} may be derived by a variation of the metric g_{ij}

 \delta g_{ij}~=~(2\delta\omega~-~\nabla_i\delta\sigma_j)g_{ij}~-~\nabla_j\delta\sigma_i 
 =~(2\delta\omega~-~\nabla_k\delta\sigma_k)g_{ij}~-~({\cal O}\delta\sigma)_{ij}, 

with ({\cal O}\delta\sigma)_{ij}~=~\nabla_i\delta\sigma_j~+~ \nabla_j\delta\sigma_i~-~\nabla_k\delta\sigma_k)g_{ij}. {\cal O} defines the conformal Killing equation ({\cal O}\delta \sigma)_{ij}u^j~=~0. Thus {\cal O} defines the number of conformal Killing vectors. The transpose of {\cal O}, {1\over 2}({\cal O}^tu)_i~=~-\nabla^ju_{ab} defines the metric moduli. The Riemann-Roch theorem states that the difference between the numbers of the metric moduli and the number of conformal Killing vectors is -3\chi, where \chi is the Euler characteristic for the manifold. This difference is then

 {1\over 2}\big(dim~ ker({\cal O})~-~dim~ ker({\cal O}^t)\big)~=~-3\chi~=~6g(g~-~1) 
{\cal O} determines the number of instanton in the theory. The size of the Goppa code is then a topological invariant given by an Euler characteristic of Floer cohomology.

This metric applies to a spatial surface with \delta g_{ij}~=~(\partial g_{ij}/\partial t)\delta t, for t a parameter labelled as "time." The Ricci flow equation

 {{dg_{ij}}\over{dt}}~=~-2(R_{ik}~+~\nabla_i\nabla_k\omega)g_{kj}, 

is an eigenvalued equation for the Ricci flow for k~=~i. The time derivative of the functional F[g,~\omega] is the low energy effective action in string theory and \omega is a dilaton field, similar to that over a string world sheet. The metric may then be rewritten according to a conformal gauge g_{ij}^\prime~=~exp(2\omega)g_{ij} and the Ricci curvature changed by
{g^{\prime}}^{1/2}R^\prime~=~g^{1/2}(R~-~2\nabla^2\omega). By an appropriate diffeomorphism, or gauge choice, the governing equation for the metric can recover the Hamilton equation for
Ricci flow.

Lawrence B. Crowell · Mar 30, 2008

The physics of the projective Hilbert space and its bearing upon spacetime physics is illuminated here by the role this plays with quantum phase transitions. The quantum critical point and the change of phase associated with it exhibits scale invariance, such as with high temperature superconductivity, and theoretically this physics is poorly understood. Current theory associates the quantum critical point with the "breakdown" of the quasi-electron and a divergence in its mass. In this critical state however the system maintains a self-similar or scale invariance.

A quantum critical state is one where fluctuations drive a phase transition at or near absolute zero temperature. To treat this sort of physics the statistical organizing principles of bosonic and fermionic states must be considered. Bosonic quantum states are by far the easiest to consider, and the correspondence between the quantum physics and the classical physics, and the emergence of different phases, are far closer and understood. A similar correspondence between fermionic quantum states and any classical physics does not exist, and the emergence of new states of matter are far less well understood. Yet in recent times the electronic states of certain transition-actinide metal crystaline systems have been measured to exist in a new scale-invariant phases, which occur at the "breakdown" of a Fermi or Landau electron liquid state and where the mass of the associated quasi-particle state diverges. This appears to be a new domain of quantum physics.

The Fubini-Study metric determines the uncertainty principle by the Berry phase as

 \phi~=~\int dt\sqrt{\langle H^2\rangle~-~\langle H\rangle^2}~=~\int dt\Delta E 

This geometric phase has the content of \Delta Et/\hbar, and for this phase bounded by \sim~2\pi[/tex] the time is then a measure of the fluctuation strength. We may of course encounter a situation, as I demonstrated in the start of this presentation, a situation where this phase may become imaginary valued. This obtains when \langle H^2\rangle~-~\langle H\rangle^2~&lt;~0. An imaginary phase may be seen as a case where the time becomes an imaginary time with t~=~\hbar/kT. For a high temperature system this imaginary time is very small, but as temperature T~\rightarrow~0 there is then more of this imaginary time available for the system to exist in an instanton or tunnelling states where this quantum behavior is detectable. We may then consider the imaginary time as analgous to the classical temperature, where for low real temperature the imaginary time assumes is analogous to a classical temperature that heats up material in the space plus imaginary time "spacetime." Thus time in general is a complex valued vector {\vec t}~\in~C^1, where ordinary time lies on the real axis of the Argand plane and temperature on the imaginary axis. For a low enough of a standard temperature the imaginary time as an internal temperature will act to heat up the quantum states of the system and they will then exhibit a phase transition or "melting," which results in a fluid behavior that is scale invariant. We then consider the case similar to the one dimensional space considered last week, indeed on 3/27/2008. Instead of a simple one dimensional space plus time model we consider the boost space of SU(1,~1) with the connection coefficient A~=~A^1\sigma_1~+~A^2\sigma_2~+~iA^3\sigma_3, The field strength tensor is F_{ab}~=~\partial_{[b}A_{a]}~+~[A_b,~A_a], where if we consider a flat connection the comonents of the curvature tensor are F_{21}~=~[A_1,~A_2],~F_{32}~=~[A_2,~A_3],~F_{13}~=~[A_3,~A_1] where the physical fields are found with the application of \epsilon^{abc} and so the time-like field is F^3~=~\epsilon F_{12}. In the case of a moving field in more than one dimension the velocity determines an action of the form S~=~\int_{r_1}^{r_2}p_rdr The Hamilton equation {\dot r}~=~\partial H/\partial p permits this to be written as S~=~\int_{r_1}^{r_2}\int_{p_1}^P_2}\frac{dr}{\dot r}dH which permits this action to be expressed according to S~=~\int tdH~\simeq~t\Delta E, for \Delta E the Fubini-Study metric distance. We then split the action into a real and imaginary part according to a complex velocity field u^a~=~(U^\mu,~iu^5), where the fifth entry corresponds to the imaginary portion of time. The field theory is then extended to five dimensions with indicial entries a~\rightarrow~(\mu,~5). For the sake of Lorentz invariance, where in general this would impose a preferred frame from which to observe the field, the complex velocity is restricted to the complex time in C^1. The field is extended similarly to a fifth dimension as well. The Hamiltonian for the above action then determines the action {\cal L}~=~-\frac{1}{4}F_{ab}F^{ab}~-~\frac{\mu^2}{2\hbar^2}u^au^bg^{cd}F_{ac}F_{bd}, where \mu has units of energy (mass). What has been done is to generalize the non-compact group into a composition of the two according to u:A_3~|\rightarrow~(A_3^\prime,~iA_3^{\prime\prime}),~A_3^\prime~\in~SU(1,1),~A_3^{\prime\prime}~\in~SU(2) or so that the complex elements are generalizations of the boosts in SU(1,1)\oplus SU(2). This means that the time-like potentials and fields have a probability of being Lorentzian or Euclidean. The above Lagrangian gives the dynamical equations \partial_aF^{ab}~=~\frac{\mu^2}{\hbar^2}(u_cu^b\partial_aF^{ca}~-~u^cu^a\partial_aF^{cb}) which when broken into its component parts is \partial_\mu F^{\mu 5}~=~0,~\partial_\mu F^{\mu\nu}~=~-(1~+~\beta^2)\partial_5 F^{5\nu}, for \beta~=~\mu u/\hbar. The term \partial_5 F^{5\nu} is a source for the field and the term \beta is then a term which amplifies the source. In the case of gravitation, which this is a "B-F" gauge like theory of, this source is the mass. The connection terms can be written according to basis elements of the manifold A^\mu~=~\partial_\nu g^{\mu\nu}, which for a bimetric theory in a reduced metric form or the traceless transverse (tt) terms for weak gravity fields gives the gravity equation \square{\bar h}_{\mu\nu}~=~\beta^2\partial^2_5{\bar h}_{\mu\nu} where the righthand side is the source G\rho and the left hand side is the Laplacian of the potential. This equation is then reduced to the Poisson equation form of Newton's law of gravity. It is interesting to compute the dispersion relationship in the gauge A_5~=~0 the spatial connection with the Fourier expansion A^\mu~=~A_0^\mu e^{ik_\mu x^\nu~+~ik_5 x^5} determines the dispersion relationships with the above dynamical equations as k_5k_\mu A_0^\mu~=~0,~\big(k_\nu k^\nu~+~(1~+~\beta^2)k_5^2\big)A^\mu~-~k^\mu k_\nu A^\nu~=~0 One of the motivating physical ideas here is with the quantum critical point for electrons in solids. The Dirac equation for the above complex time is {\cal L}_d~=~i{\bar\psi}\gamma^a\partial_a\psi~-~m{\bar\psi}\psi~-~i\frac{\mu^2}{\hbar}u^au^b{\bar\psi}\gamma_a\partial_b{\psi}, which gives the dynamical wave equation i\gamma^a\partial_a\psi~-~m\psi~-~i\frac{\mu^2}{\hbar}u^au^b\gamma_a\partial_b\psi~=~0 Again if the field is expanded in the Fourier mode \psi~\sim~e^{ik_\mu x^\nu~+~ik_5 x^5} this gives the dispersion relationship -k^ak_a~-~2\frac{\mu^2}{\hbar}(u^ak_a)^2~-~4\frac{\mu^2}{\hbar}u^au_a(u^bk_b)^2~=~m^2 which when reduced to four dimensions is -k^\mu k_\mu~=~m^2~+~(1~+~\beta^2)k_5^2. The last term for a complex time can become large and will renormalize the mass of the fermion, where as \beta~\propto~1/T will become large as the temperature approaches zero. Thus the complex time has the effect of "heating up" the internal states of the system and the mass of the fermion, in the case of the quantum critical points for the breakdown of the Landau electron liquid making the quasi-fermions "heavy." This then links aspects of a fermionic system and its quantum critical point with the instanton states of gravitation. This behavior should have a univerality to it, where it applies for all quantum fields from scalars, spinors, vectors, Rarita-Schwinger, and spin = 2 gravitons. It is also a property which determines a \beta renormalization term that is invariant with respect to scale.

Lawrence B. Crowell · Apr 13, 2008

A wormhole in spacetime is theoretically understood to result in some strange physics, such as time machines. Wormholes for a number of reasons probably don't exist as classical spacetime structures, but their role in quantum gravity is probably equal to that of quantum black holes. The results of Ford and Roman as well as the apparent logical contradictions implied by closed timelike curves likely means they don't exist as classical structures. Other authors such as Thorne and Visser disagree with this. However, this is a look at wormholes as an aspect of quantum gravity and the issue of their classical existence is not directly addressed.

The Skyrme model of QCD defines a knot topology to gluon fluxes, which in the infrared domain or a confined vacuum are scale invariant. These flux tubes wrap around each other like snakes in a mating ball. If we were to take the quarks and pull them apart with "tweezers" the flux tubes would define a braid group. I could write this in a more complete mathematical form, but I will spare myself that effort for now. So what does this have to do with quantum gravity? Conformal gravity is in the Euclidean form SU(4) and we can embed the SU(3) of QCD in this group. Then conformal gravity with the coupling constant /alpha^2~\sim~G~\rightarrow unity at unification appears remarkably QCD-ish with an extra group rank or dimension. We then consider the analogue between a quark-gluon plasma and a multiply connected spin-net. Each of the spinor fields, or spin-flags, on the manifold are analogous to quarks. In a quark-gluon plasam each quark is connected to each other through a tangle of gluon tubes which are braided and in a nontrivial topology. In the conformal gravity analogue the system each spin is connected to each other by noncommutative terms. We may think of the manifold as being inherently noncommutative, such as obeying an SU(3) group, and that there is an auxilliary U(1) gauge group which "carries" the connection terms. These is a matter of additional roots, but we will ignore that matter for the moment.

The group is then of course made Lorentzian SU(4)~\sim~spin(6)~\rightarrow~spin(5,1). The analogue of the gluon flux tubes are then multiply connected topologies between spin-flags, which are for no better term worm holes. Then as the universe expands and the coupling constant of gravity becomes weak these are removed. The manifold becomes simply connected and the connection between the spin-flags is simple. The "dual field," and in this picture QCD has a duality with gravity, the coupling constant remains large at low energy (infrared confinement) and the field theory is also Euclidean. So the quark-gluon plasma or confinement "bag" is a vacuum region with a knot topology. The braided structure of gluons are then a dual structure which are a quantum foam.

This of course is not the end of the story, for these need to be embedded in an SO(8) \rightarrow_{Z_2}~spin(7,1) or spin(8) which are a Clifford valued algebra Cl(8). The two parts are extended into F_4 and G_2 and this gives the 248 elements of the 256 N = 8 SUSY, and is the heterotic group E_8. Another approach is to use the face that E_8 containes spin(16)~\sim~ SO(32), which is a representation for the closed string. For the spin group decomposition of these heterotic groups consistently contain the DeSitter group. The spin(16) group has 128 generators. The additional 112 roots (from the total 248 in E_8) define a D_8 group (in terms of root system not lattice), which is an acceptable gauge theoretic model SO(8), which also contains the deSitter SO(3,2) under suitable change of signature. Similarly E_6 and E_7 sit inside E_8. E_6 \times SU(3)/(Z/3Z) and E_7\times SU(2)/(Z/2Z) are maximal subgroups of E_8. The simple heterotic group E_6 may be decomposed into SU(6)\times SU(2), where SU(6) contains the conformal group for gravity SU(4) by

 SU(4)\times SU(2)\times U(1)~\subset~SU(6), 

which is remarkably similar to the decomposition of SU(5) in the minimal GUT decomposition into the standard model.

With the group theory preliminaries out of the way we return to physical concepts. Consider a wormhole with openings in different regions of spacetime with different gravitational potentials. The vacua at the two openings are related by a Bogoliubov transformation between their respective raising and lowering field operators. Yet for the observer which traverses the openings the vacuum structure either abruptly changes. However since the observer in a local inertial frame carries their vacuum with then there is then a mismatch between the vacuum the observer carries in their local frame and the vacuum for a coincident local observer in that region. The multiply connected topology of this spacetime has an effective branch cut for the complex valued amplitudes of a quantum field which cycle through the wormhole. This suggests that the openings of a wormhole, or a quantum wormhole define a source for fields. Further, as the openings are multiply connected this is similar to a topologically distinct flux tube between sources. In a strong conformal gravity domain these flux tubes in SU(2) are similar to gluon flux tubes in SU(3) QCD, which connect points in the conformal spacetime to each other through "extra-dimensions. The multiply connected "tube" between the two openings then takes place in extra dimensions, such as the string/brane bulk, or for a LQG spin-net embedded in a larger space.

This then suggests that there are two principles at work here. The source, or vacua ambiguity, means that in a region of spacetime on a small enough scale will exhibit this as an uncertainty principle intrinsic to the geometry of the spacetime. This then suggests that the quantum geometry is noncommutative according to gauge potentials in a region of space or spacetime. Further, this noncommutivity is determined by a gauge potential which defines the multiply connected topology of the spacetime, similar to gluon flux tubes. In what follows these two theoretical observations are brought together to define a new approach to quantum gravity.

We set up the construction as follows. For the group SU(4)\times SU(2) we have the gauge connection A^\mu, which is an SU(2) connection, and the basis vectors e_a,~a~=~1,~2,~,3 for the space of SU(2) and \mu~=~1,~2,~\dots,~6 for SU(4). We may then impose a gauge condition on the system as

 D_\mu e_a~=~(\partial_\mu~+~{\vec A}_\mu\times)e_a. 

The gauge vector the carries an added index a, and acts as an internal vector space on the SU(4). The vector potential contracted with the basis vectors A^\mu~=~e^a A^\mu_a is a chromo-electric potential for a color charged SU(2) theory. The SU(2) basis vectors defines a bundle section and the quantum vacuum state for the according to the pure gauge element {\bf e}\wedge d{\bf e} by

 \Omega^k_\mu~=~{\epsilon_{ij}}^ke_i\partial_\mu e_j~=~B^k_\mu e_k 

The vacuum potential term B^k_\mu then determines the vacuum fields \underline\underline\Omega~=~d\inderline\Omega according to

 \Omega_{\mu\nu}~=~\partial_{\nu}\Omega_{\mu}~=~\partial_{\mu}\Omega_{\nu}~+~g\Omega_\nu\wedge\Omega_\mu~=~\big(\partial_\nu B^k_\mu~-~\partial_\mu B^k_\nu~+~g{\epsilon_{ij}}^kB^i_\nu B^j_\mu\big)e_k 

With the definition of the gauge potential A^\mu~=~e^a A^\mu_a the gauge potential

 {A^\prime}^k_\mu~=~A^k_\mu\cdot e_a~+~{\epsilon_{ij}}^ke_i\partial_\mu e_j~=~(A^k_\mu~+~B^k_\mu)e_k 

for a gauge condition with zero {A^\prime}^k_\mu we clearly have the SU(4) gauge potential defined as A_\mu~=~-B^k_\mu e_k.

Lawrence B. Crowell · Apr 20, 2008

This is a continuation of the last post

With this definition of the gauge potential A_\muand the vacuum potential \Omega_\mu we examine gauge transformations by the infinitesimal change in the sectional basis \delta {\vec e}~=~{\vec\epsilon}\times{\vec e}, for \vec\epsilon an infinitesimal vector displacement of vectors in S^3 of SU(2). The vector symbol is for the vector part on S^3 For a general {A^\prime}^k_\mu determined from A^k_\mu~=~-B^k_\mu by a gauge covariant gauge displacement vector \chi^k_\mu give the transformation

 e_i\delta A^i_\mu~=~e_i\partial_\mu\epsilon^i,~e_i\delta{A^\prime}^i_\mu~=~D_\mu e^i 

 \delta B^i_\mu~=~-\delta B^i_\mu,~\delta\chi^i_\mu~=~-(\epsilon\times\chi_\mu)^i 

which illustrates that the covariant gauge vector \chi^i_\mu and the gauge potential are gauge independent. This covariant gauge vector is a chromodynamic-like field similar to a gluon, or the weak gauge bosons. These potentials determine the physical fields

 F^i_{\mu\nu}~=~\partial_\nu A^i_\mu~-~\partial_\mu A^i_\nu,~e_iG^i_{\mu\nu}~=~-{\vec e}\cdot\partial_\nu{\vec e}~-~{\vec e}\cdot\partial_\mu{\vec e}~=~e_i(\partial_\nu B^i_\mu~-~\partial_\mu B^i_\nu) 

 {F^\prime}^i_{\mu\nu}~=~F^i_{\mu\nu}~+~G^i_{\mu\nu} 

This then indicate that the vector portions of gauge connections A^i_\mu transform identically with the SU(2) portion of the connection B^i_\mu.

The basis vector \vec e defines a phase U~=~e^{\phi\sigma\cdot\vec e}, for \sigma a vectors of Pauli matrices. A momentum for a particle in the SU(4) space is p_\mu~=~U\partial_\mu U, which when defined in a basis with cos(\phi)~=~0 gives the Lagrangian

 {\cal L}~=~\frac{1}{2}p_\mu p^\mu~+~\frac{1}{16}Tr\big([p_\mu,~p_\nu]\big)^2 
 =~-\frac{1}{2}G^\prime_{\mu\nu}G^{\mu\nu}~-~B_\mu B^\mu, 

for G^\prime_{\mu\nu}~=~G_{\mu\nu}~+~B_\mu\wedge B_\nu. This Skymre Lagrangian then leads to the dynamical equation

 {\vec e}\times\nabla^2 {\vec e}~-~(\partial_\mu G^{\mu\nu})\partial_\nu{\vec e}~=~~0 

This differential equation is \partial^\mu j^i_\mu~=~0 for the term

 j^i_\mu~=~({\vec e}\times \partial_\mu {\vec e})^i~-~G_{\mu\nu}e^i, 

where the continuity equation defines a conservation of charge across flux tubes. With ({\vec e}\times\partial_\mu{\vec e})^i~=~\epsilon^{ijk}B^k_\mu e^j the current components are clearly of the form

 j^1_\mu~=~B^1_\mu~+~{G_\mu}^\nu}(B^2_\nu e^3~-~B^3_\nu e^2),~j^2_\mu~=~B^2_\mu~+~{G_\mu}^\nu}(B^3_\nu e^1~-~B^1_\nu e^3),~j^3_\mu~=~B^3_\mu~+~{G_\mu}^\nu}(B^1_\nu e^2~-~B^2_\nu e^1) 

The third of these equations may be removed by the gauge condition D_\mu e^3~=~0, which reduces the problem to two dimensions. The two continuity equations found from \partial_\mu{\vec j}^\mu~=~0 are then

 \partial^\mu B^1_\mu~+~B^{3\mu}B^2_\mu~+~\partial^\mu{G_\mu}^\nu B^2_\nu~=~0 
 \partial^\mu B^2_\mu~-~B^{3\mu}B^1_\mu~-~\partial^\mu{G_\mu}^\nu B^1_\nu~=~0 

This equation may be replaced by the simple substitution \omega^{\pm}_\mu~=~(B^1_\mu~\pm~iB^2_\mu)/\sqrt{2}, which defines creation and annihilation operators for the SU(2) field. The differential equation may be written as

 \omega^{\pm}\partial^\mu\omega^{\pm}_\mu~\mp~\omega^{\pm}B^{3\mu}\omega^\mp_\mu~=~\pm\omega^{\pm}\partial^\mu{G_\mu}^\nu \omega^{\mp}_\nu 

which when integrated over S^3 the right hand side defines a charge according to a Chern Simons index

 kQ~=~\pm\int d^3x\omega^{\pm}\partial^\mu{G_\mu}^\nu \omega^{\mp}_\nu~=~\pm\int d^3x\epsilon_{ijk}\epsilon^{\mu\nu\sigma}\omega^i_\mu\omega^j_\nu\omega^k_\sigma 

which is also an index for the knot equation. This index is then a topological invariant for the knot topology of the "gluon-like" threads.

This gauge theory with gluon threads on the manifold for conformal gravity connects spin-gravity terms in a multiply connected manner, which might be interpreted as wormholes. The index derived defines a topological invariant for this multiple connectivity in conformal gravity embedded in SU(4)\times SU(2). This structure will next be used to derive a form of noncommutative quantum gravity embedded in the heterotic group E_6, which in turn is embedded in E_8.

Lawrence B. Crowell · Apr 28, 2008

The gluon-like threads on the conformal space is define alternative assignments for how points are connected to each other. One connection between points in conformal gravity might be a path on the spin(4,2) manifold of conformal gravity, while another connection is along the gluon-like flux tube. This may be considered from the perspective of a thickened S^2~\subset~S^3, where the knot equation is defined. The S^2 is a bloch sphere defined by the gauge condition, where every point on this sphere is associated with a copy of spin(4,2). The Cartesian product of conformal gravity with the isposin group defines a fibration on the two sphere by spin(4,2), every bit as much as a spinor fibration on the manifold of conformal gravity. Thus for the spinor field connection A_\mu^i[/itex] the multiple connectivity between two points may be given by the covariant x^\mu~=~y^\mu~+~\Omega_i^{\mu\nu} A^i_\nu The spinor group defines a set of noncommutative coordinates y^\muon the space of conformal gravity due to the internal rotation on the Bloch sphere at each point on the conformal manifold [y^\mu,~y^\nu]~=~i\Omega_i^{\mu\nu}e^i where \Omega_i^{\mu\nu}e^i[/itex] is a symplectic matrix. The orientation of each Bloch sphere on the conformal manifold will deviate by the curvature on that manifold, as well as due to an internal gauge rotation. A general field over the set of noncommutative coordinates established by local interal rotations is then \phi~\in~{\cal A}_\omega. With an intertwiner or n-trad basis E^i_\mu the coodinates on a Bloch sphere may be defined as y^i~=~E^i_\mu y^\mu so that a covariant set of coordinates on the spinor space is &lt;br /&gt; x^i~=~y^i~+~\Omega^{ij}_\mu A^\mu_j.&lt;br /&gt; Either perspective is equally valid. This n-trad basis is then a frame where on spin(4,2) for a given local basis on the Bloch sphere, or may equivalently define a frame on the Bloch sphere for a covariantly constant basis on spin(4,2). The noncommutative scalar field induced by internal spinor rotation defines a two dimensional space determined by a local patch on each Bloch sphere at every point on the conformal manifold. This space is then a tesselated space R^2~=~{\bf U}_i^{\infty} N_i, where each N_i is a neighborhood on each S^2 around where two state vectors point in adjacent or neighboring Bloch spheres. This is a form of an atlas-chart construction which can be explicitely constructed. The transition function between the two charts, or Bloch wheres, defines on R^2 a space of connection terms, which is {\cal A}_\omega. For a general curved space, here the conformal space, the \sigma_i directions will vary and thus define at each point on the spin(4,2) a unique eigen-basis. This infinite set of of eigen-bases then is what the field is expanded upon in an orthogonal and complete set of states &lt;br /&gt; \phi(y)~=~\sum_{m,n}|m\rangle\langle n|.&lt;br /&gt; This defines {\cal A}_\omega as a C^* algebra who&#039;s generators act on an element \phi(y)~\in~{\cal A}_\omega according to &lt;br /&gt; e^{ik\cdot y}\phi(y)e^{-ik\cdot y}~=~\phi(y~+~k\cdot\Omega)&lt;br /&gt; The coordinates y_\mu define the adjoint action of the spinor group on the spin(4,2) manifold according to &lt;br /&gt; ad_{x_\mu}[\phi]~=~[x_\mu,~\phi(y)]~=~i\Omega^{ij}\frac{\partial x_\mu}{\partial y^i}\frac{\partial\phi}{\partial y^j},&lt;br /&gt; where V^i_\mu \partial_i~=~\Omega^{ij}\frac{\partial x_\mu}{\partial y^i}\frac{\partial_j} is a gravitational vielbein for a spacetime metric. The intertwiner E^i_\mu between the two spaces defined by SU(4)~\sim~spin(4,2) and SU(2) defines the simplectic form \Omega_{\mu\nu}, which gives the inverse matrix \langle B_{\mu\nu}\rangle~=~(\Omega^{-1})_{\mu\nu}, which is the vacuum state for the gauge fields, as indicated above. The vielbein above also has its inverse analog U^\mu_i \partial_\mu~=~\Omega^{\mu\nu}\frac{\partial x_i}{\partial y^\mu}\frac{\partial_\nu} The space R^2 is defined by the difference between state vectors on Bloch spheres within a chart. The Bloch sphere is defined by Hilbert state vectors modulo the phase, which is the projective Hilbert space PH. Hence the spaces SU(4) and SU(2) represent a double fibration over the projective Hilbert state space. Each these spaces defines a fibration on the projective Hilbert space. The general relativity portion of the conformal group is then a fibration on the state space which is a more complete version of the above discussion on Hawking radiation from an elementary consideration of the Berry phase. The spinor group determines another fibration on the projective Hilbert space, which determine currents or flux tubes between different point on the conformal manifold. Alternatively it defines noncommutative coordinates on the conformal space, which are the gauge parameterization of the quantization of the conformal gravity theory. This double fibration with the sympletic matrix \Omega when it acts on a basis vector \partial_{x} on either the manifold of conformal gravity or the space of SU(2) defines the sympletic dual as p~=~\omega\cdot\partial_{x}. Thus the double fibration is analogous to the one-form \omega~=~\gamma(dx~-~vdt), which can be the Lorentz boosts, or a Lagrangian form used in Finsler geometry. Thus the spaces SU(4) and SU(2) define the two fibrations over the projective Hilbert space according to the vacuum state, equivalently the symplectic form, for the field which defines a noncommuting basis over the spacetime. The term ad_{x_\mu}[\phi] defines the symplectic variable conjugate to x^\mu, which describe a quantum fluctuation due to noncommutative variables. In what will follow the noncommutative spacetime is then a coarse grained perspective on the manifold, which is a quantum phase transition. For the full group E_6, or even further E_8 the noncommutative basis transitions into a full heterotic gauge group action that has no scale dependency. But then at the symmetry breaking E_6~\rightarrow~SU(4)\times SU(2)\times U(1) the universal scale invariance of the heterotic group is lost. In the full heterotic group E_8 there exist 16 scalar fields 4\times(2~+~{\bar 2}), which are dilaton/Higgsian type of fields, of which 9 are associated with the conformal SU(4) and the other six are associated with the breaking of the full heterotic E_8.

Lawrence B. Crowell · May 4, 2008

Gravitation and gauge fields, such as in the standard model, exhibit a different invariance with respect to vacuum energy rescaling. The Lagrangian for general relativity is L~=~R~+~2\Lambda, for R the Ricci scalar. Thus under a rescaling of the cosmological constant the momentum energy tensor is changed by T_{ab}~\rightarrow~T_{ab}~-~(c^4/8\pi G)\Lambda g_{ab}. This then breaks the symmetry of the gauge field Lagrangian invariance under vacuum rescaling. The cosmological constant then introduces the global curvatures of a cosmology as an Einstein space directly into the gauge field, as well as the Dirac Lagrangians for the source fields or particles. The cosmological constant, with units of inverse length squared and the coupling constant (c^4/8\pi G)~=~(\hbar c/8\pi)L_p^2 give a unit of energy density

 \rho~=~\frac{c^4}{8\pi G}\Lambda~=~(L_pL_c)^{-2}, 

where L_cis a length associated with the cosmological constant L_c~=~\sqrt{1/\Lambda}~\sim~10^{27}cm, or about the distance to the cosmological event horizon. This then means the vacuum energy density would be approximately \rho~\sim~10^{12}cm^{-4}, or approximately 10^{-46}GeV^4. This vacuum energy connects two scales, one which is the smallest scale whereby physics may well breakdown on smaller scales, and the second a cosmological scale, approximately where the cosmological event horizon lies r~=~\sqrt{3/\Lambda}. On scales larger than this cosmological scale quantum amplitudes across the cosmological event horizon are related by Bogoliubov transformations and the vacuum states for fields separated by this distance are not equivalent. Further, on this cosmological scale the homogeneity of time breaks down. As the metric components for the deSitter cosmology explicitely depend on time, such as g_{rr}~=~exp(\sqrt{\Lambda/3}t), there is no way in which there can exist a Killing vector which acts to maintain energy, or the four components of the momentum four vector, holonomically constant. Consequently the constraints on physics according to the Coleman-Mandula theorem, where these symmetries permitted conserve energy, are locally restricted. They pertain within a scale (r~&amp;gt;~L_p,~r~&amp;lt;&amp;lt;~L_c~=~10^{60})L_p, where vacua structure separated by distances approaching the cosmological distance L_c become unitarily inequivalent. These two scales L_p,~L_care comparable during the big bang. The early high energy universe, or on a scale where E_8 has been broken and spacetime was governed by SU(4)\times SU(2)\times U(1). The gauge induced wormholes, governed by SU(2), act strongly on the early universe by the shifting T_{ab}~=~T_{ab}~-~\Lambda g_{ab}, where the cosmological constant was large during this inflationary period. As one approaches t~\rightarrow~0 the two scales close L_c~\rightarrow~L_p. Under these conditions the classical spacetime structure fragments into a noncommutative froth as unitarily inequivalent vacua are compressed together into ever small spacetime volumes which become classically indefinable. The symplectic matrix \Omega_{\mu\nu} defines the vacuum energy density in any local region as &lt;br /&gt; \rho~\simeq~|\Omega_{\mu\nu}|^{-2}~=~|\langle G_{\mu\nu}\rangle|^2~\sim~\frac{1}{(L_pL_c)^2},&lt;br /&gt; where G_{\mu\nu} is the field strength tensor of the SU(2) gauge field. The expansion of the spacetime under the inflationary pressure of a large cosmological constant reduces the influence of the gauge field on the conformal spacetime on larger scales. The very early universe after the breaking of the heterotic symmetry is defined as a system of gravity-like excitons which have multiple connections in a spin net and further are defined on vacua which are not unitarily equivalent. The expansion of the universe and the establishment of a stable cosmological horizon length by the Higgsian vacuum extends regions of unitary equivalence over vast distances \sim~10^{10} light years. As the universe expands and ages it asymptotically approaches a deSitter spacetime configuration with a diminishing mass-energy content. Eventually Hawking-like radiation from the cosmological horizon will very slowly cause the cosmological horizon radius increase to infinity as the spacetime approaches a Minkowski spacetime as {itex]t~\rightarrow~\infty. This final state of the universe is the AdS conformal infinity, which is shared by the dS cosmology, since (SO(4,1),~SO(3,2))~\subset~SU(4). The evolution of the universe is then a process which maps a set of inequivalent vacua, a purely quantum system of excitons, into a completely classical spacetime configuration with \rho~=~0. The universe is then a map between these voids. The complete symmetry of the universe is then some form of quantum error correction code which preserves the total quantum information through the process. The quantum error correction code is then a Golay or Goppa code, such as that defined by the Leech lattice \Lambda_{24}, which includes three E_8 heterotic groups in a modular system.

Lawrence B. Crowell · May 11, 2008

I have so far laid down an intermediate energy hypothesis for quantum gravity which stems from the breakdown of the heterotic group E_6 into a spinorial gauge-like field plus conformal gravity plus an abelian group. The spinorial gauge-like field is SU(2) and induces a noncommutative spacetime basis on conformal gravity that is a spin-net. The non-commutative aspects of the spin-net then describe a holonomy for any enclosed path. The structure of this holonomy is comparatively simple, as seen in Wheeler's trick of rotating a book tied to a ribbon with the other end fixed twice or by 4\pi radian rotation. In general the parallel translation of a vector along a path connecting the two points (x_0,~y_0,~z_0) to (x_1,~y_1,~z_1) rotates that vector around the axis \vec a

 {\vec a}~=~\kappa {\vec x}_0\times{\vec x_1}, [/itex] where \kappa is a curvatures scalar related to the Gauss' fundamental forms. The parallel translation around a small parallelogram of area A means that for small angle deviation and {\vec x}_0\times{\vec x_1}~=~|x_0||x_1|sin(\theta)~\simeq~|x_0||x_1|\theta that this angle is given by \theta~\simeq~2A\kappa, which is the result for the Ricci curvature scalar on a small region of constant curvature. In a similar light a spinor may be parallel translated. A particle with a spin j\hbar, where the Dirac unit of action is set to unity, is transformed by a unitary operator U_j, which is a group element of the SU(2), which is a group of homomorphisms between spin states. For two points on the base manifold we may define a curve C, or in a sufficiently local small region an edgelink, which defines the amplitude for a spin state at the start and end point of this curve as A_{C}(m_1,~m_2)~=~\langle j_1, m_1|U_j|j_2,~m_2\rangle, for m the eigenvalue of the j_z spin operator. This amplitude defines the elements of the unitary U_j, \langle j_1, m_1|U_j|j_2,~m_2\rangle~=~\langle j_1, m_1|(z|j_2,~m_2\rangle), for z a complex number. This may then be used to define the amplitude for the rotation of a state |j_3,~m_3\rangle as the output from a node n A_n(m_1,~m_2,~m_2)~=~\langle j, m_1|(z|j,~m_2\rangle\otimes|j,~m_3\rangle). This may be interpreted as the input of two spins j_1,~j_2 from two edgelinks into a node and the output state j_3. The complex number is evaluated from the unitary operators for the spins j_1,~j_2,~j_3 as A_{n}(m_1,~m_2,~m_3)~=~(U_{j_3}\odot\langle j_3,~m_3|) z(U_{j_1}|j_1,~m_1\rangle\otimes U_{j_2}|j_2,~m_2\rangle) ~=~\langle j_3,~m_3| U_{j_3}^{-1}U_{j_3} z(|j_1,~m_1\rangle\otimes |j_2,~m_2\rangle) ~=~\langle j_3,~m_3| z(|j_1,~m_1&gt;\otimes|j_2,~m_2\rangle) . Then from the individual unitary spin operators the complex number z is seen to define a map which reduces tensor products from the incoming states to the outgoing state, or equivalently z is a Clebsch-Gordon coefficient. In an interative manner these edgelinks may be combined to define a spin network, which also describes an underlying non-commutative structure to spacetime. Any edgelink between two points x,~y, c_{x,y} carries a representation of the SU(2) plus that of conformal gravity. These quantum numbers are j,~q for spin and charge, here the charge being mass. The non-commutative structure of the spacetime is due to a disordered nonlocality or multiple connectivity. We might think of the SU(2) as defining a wormhole that connects two points along another connected path. Any holonomy will is defined according to a loop with an ambiguity with respect to the ordering of points. Energetically this disordered nonlocality is a frustrated system, and can only remain within this ambiguity within the Heisenberg uncertainty in energy. Classically this ambiguity is forbidden. Up to this point the subject of supersymmetry has been avoided. Supersymmetry introduces an underlying structure to spacetime. This is a graded algebraic system that defines a supermanifold with coordinates y^\mu~=~x^\mu~+~i\theta^\alpha\sigma^\mu_{\alpha\dot\beta}{\bar\theta}^{\dot\beta}. To the classical structure of the manifold is a spinorial term due to Grassmannian generators \theta,~\bar\theta. N=2 supersymmetry permits an SU(2) symmetry to a manifold with the two supergenerators Q^1_\alpha,~Q^2_\alpha. In what follows some familiarity with supersymmetry is presumed. The N=1 Grassmannian generators are \theta^1,~{\bar\theta}^1,~\theta^2,~{\bar\theta}^2. A supermanifold coordinate is then given by y^\mu~=~x^\mu~+~i\theta_1^\alpha\sigma^\mu_{\alpha\dot\beta}{\bar\theta}_1^{\dot\beta}~+~i\theta_2^\alpha\sigma^\mu_{\alpha\dot\beta}{\bar\theta}_2^{\dot\beta} The commutator between two super-coordinates is calculated to be [y^\mu,~y^\nu]~=~\sigma^{\mu\nu}(\theta_1\theta_2 \sigma^\mu\cdot A_\mu~+~H.C.) where \{{\bar\theta}_1,~{\bar\theta}_2\}~=~\sigma^\mu\cdot A_\mu, which is a gauge potential. This is similar to the symplectic form of the non-commutative coordinates previously derived. An anti-commutator between two supergenerators Q_i^\alpha gives \{Q^i_\alpha,~{\bar Q}^j_{\bar\beta}\}~=~2\sigma^\mu_{\alpha{\bar\beta}}(\partial_\mu~+~A_\mu), [/itex] which is the gauge covariant SU(2) operator. This then links the internal gauge symmetry for the non-commutativity of spacetime is due to supersymmetry, or is a gauge action due to an N~\ge~2 supersymmetry. From here we have a link to supersymmetry types and their small groups. This will begin to connect us to the 24-cell and quantum error correction code systems.

Lawrence B. Crowell · Jun 22, 2008

For a black hole which emits a particle the back reaction g~\rightarrow~~g' is a classical convenience. It is almost a Copenhagen-esque treatment of the black hole as a classical system which collapses the decoherent state of the emitted quanta (photons etc). Yet for a small enough black hole the metrics for the two black hole configurations differ enough so that the Killing vectors K and K' can't be regarded as equivalent under any sort of "nice" map. To do so is to say that the two metrics have a coordinate to coordinate relationship which violates the covariance principle of relativity. Thus there is no diffeomorphism which can be established to tie one metric to the other --- they pertain to different spaces.

In what follows there is a fairly extensive discussion on mathematics and some conjectures. This is induced by a physical argument that the Wick rotation from a Euclidean to a Lorentzian metric induces a breaking of general covariance between the configuration metrics and decoherence. This gives rise to an energy fluctuation with an associated time uncertainty. This time uncertainty is what gives rise to a macrotime with a Lorentzian signature. Once this mathematics is presented, the existence of knot topology is argued for with a braid group structure PSL_2(Z). This leads to a connection between quantum states of gravity for metric configuration variables with the Lorentzian signature and a Fermi surface. The purpose of this here is to construct a one to one relationship between a Lorentzian spacetime structure and its Euclidean analogue. This will then permit a universal scaling of physics with imaginary time fluctuations \delta t~=~\hbar/kT.

So to first reiterate, let the manifolds (M,~g) and (M',~g') be related by a function f so that f:M~\rightarrow~M'. For these two metric homeomorphic to each other we have that

f is bijective (injective & surjective),

f is continuous,

the inverse function f^{-1} is continuous (f is an open mapping).

If such a f:M~\rightarrow~M' exists then M and M' are homeomorphic. Homeomorphisms define an equivalence relation between topological spaces, eg a coffee cup and donut are equivalent. Classes of topological spaces are homeomorphism classes.
Is this homeomorphism a diffeomorphism? A diffeomorphism is a homeomorphism that is C^r, r~\ge~1 differentiable. But as we have seen if the map between the metrics g~\rightarrow~g' is differentiable then \partial U/\partial x exists and this is (in coordinate terms)

 {{\partial U^a_b}\over{\partial x^c}}~=~{{\partial^2y^a}\over{\partial x^b \partial x^c}} 

which is a non-covariant map between metrics. Thus the homeomorphism between the two spaces is not a diffeomorphism.

Consider a set of R^4s in the Euclideanized domain, in particular for disks D^4 on a Euclideanized DeSitter spacetime, exist as a set of all possible small R^4s homeomorphic to the disk. We might then consider them as classical configurations, eg non-holomorphic outcomes (state reductions) from the wave functional for spacetime \Psi, for quantum states. In the quantum tunnelling with t~\rightarrow~it there is then a continuation of this homeomorphic structure into the Lorentzian domain of possible spaces. These spaces are similarly not identified to each other by diffeomorphisms. As such the path integral over all possible spacetime configurations is coarse grained into decoherent sets, where these sets are over fields that are diffeomorphic, but decoherence exists between topological spaces that are not diffeomorphic.

With two metrics g and g', with different Killing vectors we might find an estimate of the change in gravitational self-energy between the two as

 \delta E_G ~= |\Gamma~-~\Gamma'|^2, 

where the \Gamma,~\Gamma' are connection terms on the two metric spaces. This energy uncertainty obeys \delta E_g~\sim~\hbar/\delta T. For a weak gravity field these terms will simply be \Gamma~\sim~\nabla\Phi, where \Phi is the Newtonian gravitational potential. So we have as our approximation

 \delta E_G ~= |\nabla(\Phi~-~\Phi')|^2, 

where we might set \Phi~-~\Phi'~=~f, and consider this function in a more general setting. This energy functional on the unit ball B, or B~homeo~B', defines the map or function f:B~\rightarrow~R,

 E(f)~=~\int_B|\nabla f|^2 db 

For the functional to be continuous, the vector space of all such functions need a well behaved topology. This is then well defined in a Euclidean metric, but no the Lorentzian case since the moduli is nonHausdorff. Also for 4-manifolds it can be the case that the ball B has a homeomorphic identification with another such ball (small exotic R^4's) that is not diffeomorphic.

This leads to some conjectures on the relationship between nonHausdorff moduli space, which pertain to the Lorentzian spacetime, and strange properties of four dimensional manifolds determined by the intersection form for such manifolds. Physically this should provide a map between a superposition of states, each over a unique metric configuration variable, and a semiclassical state weighted around a classical state with a nonHausdorff moduli space. The inability to separate different moduli in the Lorentzian case has a relationship with the homeomorphic maps between configuration metrics in the Euclidean case. The terms in the moduli space are \Lambda^1(M_l)/{\cal G} and the curvature terms F~=~D\Gamma in the Euclidean case are in \Lambda^2(M_e), and the intersection form {\underline\omega}^4~=~F\wedge F exists in \Lambda^4 and 4\pi k~=~\int{\underline\omega}^4 is a Chern class.

The set of 4-manifolds, which includes these strange "fake" spaces, have their homeomorphic structure determined by the embedding to two-disks or two-spheres in the space. In one dimension higher this structure leads to a knot topology. Whitney's theorem tells us we can embed a 2-dim space into a 5 dim space. We have a curious ambiguity about embedding 2-dim spaces in 4-dim spaces, for neither of Whitney's theorems apply. Yet we have all this duality/self-duality machinery, which suggests there is some embedding property. Suppose that we have a five dimensional space as R^4\times[0,~\epsilon], where \epsilon is some small number. So we have a thickened R^4. Now consider a disk D^2, where maybe there are "curvatures" from F_{ab} or the Weyl curvature C^{ab}_{cd}. This cap might have been removed from some 2-dim space (call it X^2) we embed in our R^4\times[0,~\epsilon]. If this cap is a cross-cap (points identified with an x~\rightarrow~-x for nonorientable structure), or with some double covering or ..., this then contains some topological information of X^2. We may then isolate all non-trivial topology of X^2 on a set of disks or caps. We know we can embed this D^2 in this 5-dim space by Whitney's theorem. Further, the dual curvatures *F_{ab} or *C^{ab}_{cd} for * = Hodge dual operator, will exist on a dual cap or disk *D^2 if we restrict to R^4 or assume curvature two-forms with direction in the 5-direction are approximately zero. Again Whitney tells us that we can embed this in our thickened R^4.

A theorem by Taubes tells us we can "collect" all the topological charge on a manifold given by the intersection form into a finite number of points and maintain the topology of the space. So as a result I am going to propose that all of the relevant curvature information is restricted to these disks with all of the curvature information. So we have a set of disks \{D^2_i,~ i~=~1,~\dots,~n\} and their dual disks \{*D^2_i,~ i~=~1,~\dots,~n\}, with curvature information and respective duals.

Now for curvatures given be a connection one-form A we may evaluate on a disk *D^2 the Wilson line

 W(C)~=~exp(i \int_C A), 

for the curve C on our disk D^2. The expectation of this Wilson curve is the partition function or path integral

 \langle W(C)\rangle~=~\int D[A] W(C)exp(-iS), 

where S is the action determined by the curvature two-form F~=~DA. Thus we have there is the skein relationship for over-crossings and under-crossings of lines with

 \alpha\langle W(L^+)\langle~-~\alpha^{-1}\langle W(L^-)\rangle~=~z\langle W(L^0)\rangle 

\alpha~=~1~-~2\pi i/kN, z~=~-2\pi i/k N, and L^+ the over-crossing lines, L^- the undercrossing lines and L^0 the noncrossing lines

This suggests that the Whitehead torsion is related to knot polynomials or Braid groups. In other words the Whitney trick works in 5-dim, and it appears that the cobordism is related to the knot polynomial or braid group, or that the Whitehead group W(M)~=~B_n or W(B_n), here restricted to braid groups. For n~=~2 the braid group is the cyclic group, and W(Z) is trivial. For n~>~2 where

 s_i s_j~=~s_j s_i [/itex] s_i s_{i+1} s_i~=~s_{i+1}s_i s_{i+1}~Yang-Baxter~equation [/itex] the braid group is the trefoil knot group for n~=~3 and for n~&amp;gt;~2 these groups are infinite nonabelian groups. It appears that in general W(B_n) is trivial. These crossings might be considered to be punctures through the disk D^2, and as such have homotopy content. The permutation of these punctures, given by the Braid group relationship, is a homotopy. This makes some contact with the idea that W(M)~=~W(B_n), here in the case where we are looking at just one disk on the X^2. This appears to be a way of looking at what happens with R^4~=~\lim_{e--&amp;gt;0}R^4\times [0,~\epsilon]. The modular braid group is PSL_2(Z), which is an orbifold configuration on a hyperbolic group SL_2(C)~=~SU(2)\times SU(1,~1), or SO(2,1). This is the group for anyons. The first homotopy group of SO(2,1) \pi_1(SO(2,1)~=~Z, or infinite cyclic. The braid group is a projective representation of the special orthogonal group SO(2,1), which are representations of anyons. For later I will discuss the structure of the Fermi surface. This will connect with my prior discussion of a universal scaling principle with spin fields off a Fermi surface which give a Landau electron-like liquid.

Lawrence B. Crowell · Sep 7, 2008

This is a note on how a Bloch wave analysis of quantum states and fields may be used in quantum field theory and gravitation. This is an approach which uses some results of solid state physics. There have been some approaches of this nature in lattice gauge field theory and there have been some connections with ion traps and quantum gases.

The wave function in a crystalline solid will have a periodicity which matches the occurrence of ions in the lattice. The lattice defines a translation operator T({\vec r})~=~exp({\vec k}\cdot{\vec r} with a periodicity relationship T({\vec r}+{\vec a})~=~T({\vec r}), for |\vec a| the lattice distance. The wave function for a particle is then translated by this operator as

 T({\vec r})\psi({\vec x})~=~\psi({\vec x}+{\vec r}). 

The wave function also will exhibit local phase shifts according to a gauge group \cal G. The wave function transforms according to {\cal G}:\psi(\vec x)~\rightarrow~\psi({\vec x})e^{i\theta}. The action of the gauge group and the translation operator is to change the wave function by

 \psi({\vec x})~\rightarrow~UT(\vec r)U^\dagger\psi(\vec x)~=~\psi({\vec x}+{\vec r})e^{i\theta}. 

Consider the wave function \psi({\vec x})~=~\psi_0exp(-i{\vec k}\cdot{\vec x}). The action of the group operators U and the translation operator T(\vec r) is then to shift the position variable of the wave function in the position representation. The combined action of the group operator and the translation operator for small group actions U~\simeq~ 1~+~i\theta and small translations T({\vec r})~\simeq~1~+~i{\vec k}\cdot{\vec r} is

 UT(\vec r)U^\dagger~\simeq~(1~+~i\theta)(1~+~i{\vec k}\cdot{\vec r})(1~-~i\theta) 

 =~1~-~[\theta,~{\vec k}\cdot{\vec r}]~=~1~-~{\vec k}\cdot [\theta,~{\vec r}]~-~[\theta,~{\vec k}]\cdot{\vec r}~+~O(\theta^2). 

The action of this composite operator is to translate wave function from \vec x to \vec x' so that x'~=~x~-~[\theta,~{\vec r}], which maybe written according to

 x'~=~x~+~iU\nabla_kU^\dagger. 

In the conjugate momentum representation for T~=~T(\vec k) which translates the wave function on the reciprocal Brillioun lattice there is the conjugate result that

 p'~=~p~+~iU\nabla_rU^\dagger. 

The directional derivatives of the group elements define potential terms {\vec A}~=~U\nabla_r U^\dagger and the conjugate potential \vec C~=~U\nabla_kU^\dagger, where in the first case this gives the standard gauge covariant definition of the momentum operator. This results in the O(\hbar) commutation relationships

 [x,~x']~=~i\nabla_{[k}C_{k']},~[p,~p']~=~i\nabla_{[r}A_{r']},~[x,~p]~=~i~+~i\nabla_{r}C_{p}~-~\nabla_{p}A_{r}, 

which is a quantum system with noncommutative coordinates. These noncommutative variables are defined by gauge potentials which are evaluated around a loops in position and momentum space, which are a form of Berry phase. The potential terms A_r and C_k, which we can in general write as A_z defines a quantum magnetic-like field B_{z'}~=~(\nabla_{[z}A_{z'']})_{z'}, which has a clear analogue with the Maxwellian result {\vec B}~=~\nabla\times{\vec A}. This quantal magnetic field is entirely due to the lattice structure of the space, which can be a lattice in solid state physics or a winding number for string on orbifolds. In the case of solid state physics this is related to quantum hall effects and magnetic flux quanta in periodic solids.

The Dirac equation is the square root of the Klein-Gordon equation and constitutes a quaterionic system. The Dirac algebra is a 32-part algebra produced with the combination of the four-vector units (i,~{\bf i},~{\bf j},~{\bf k}) with the unit quaternions (1,~{\bf e}_1,~{\bf e}_2,~{\bf e}_3). Physically the two sets denote spacetime and mass-charge parameters. These enter into the dynamics of a relativistic particle with the relativistic momentum-energy invariant interval

 m^2~=~E^2~-~p^2. 

This may then be factorized according to the quaternion elements with the additional phase term e^{-i(et~-~{\bf p}\cdot{\bf x})}, so this interval assumes the form

 (\pm {\bf e}_3E~\pm~i{\bf e}_1{\bf p}~+~i{\bf e}_2 m)(\pm {\bf e}_3E~\pm~i{\bf e}_1{\bf p}~+~i{\bf e}_2 m)e^{-i(et~-~{\bf p}\cdot{\bf x})}~=~0. 

By the standard quantization rule E~\rightarrow~i\partial/\partial t and {\bf p}~\rightarrow~-i\nabla the energy momentum interval is reproduced as

 \Big(\pm i{\bf e}_3{\partial\over{\partial t}}~\pm~{\bf e}_1\nabla~+~i{\bf e}_2m\Big)(\pm {\bf e}_3E~\pm~i{\bf e}_1{\bf p}~+~i{\bf e}_2 m)e^{-i(et~-~{\bf p}\cdot{\bf x})}~=~0. 

The wave function is then a quaternionic state vector

 \psi~=~(\pm {\bf e}_3E~\pm~i{\bf e}_1{\bf p}~+~i{\bf e}_2 m)e^{-i(Et~-~{\bf p}\cdot{\bf x})}~=~0, 

which obeys the invariant interval is the quantum wave equation

 \Big(\pm i{\bf e}_3{\partial\over{\partial t}}~\pm~{\bf e}_1\nabla~+~i{\bf e}_2m\Big)\psi~=~0. 

where the basis vectors define the Dirac matrices. It is clear that for a massless system that for the right hand side with the phase term boosted that the Dirac equation is equivalent to

 (\pm {\bf e}_3E~\pm~i{\bf e}_1{\bf p})(\pm {\bf e}_3E'~\pm~i{\bf e}_1{\bf p}')e^{-i(Et~-~{\bf p}\cdot{\bf x})}~=~0, 

for {\bf p}'~=~{\bf p}~+~{\bf A}_r. The Dirac equation then predicts that the momentum vector and energy evolves according to

 {{\partial {\bf p}}\over{\partial t}}~+~\nabla_r E~+~[E,~{\bf p}']~=~0 

 {{\partial E}\over{\partial t}}~-~\nabla_r {\bf p}~+~[{\bf p},~{\bf p}']~=~0 

The first of these differential equations is a modified Hamilton equation due to the addition of a quantum correction term, and the second term is a noncommutative or quantum corrected version of the invariant interval m^2~=~E^2~-~p^2. The commutator is determined by the action of the translation operator upon the group elements. In the case where the group contains the Lorentzian group the gauge potential will in general be A_r~=~A_r~+~\omega\wedge A_r, for \omega the gravitational connection term. The quantum correction due to the noncommutative coordinates will then be of the form

 [{\bf p},~{\bf p}']_\mu~=~(\nabla_{[r}A_{r']})_\mu~+~R_{\mu\nu\alpha\beta}U^\nu p^\alpha p^\beta 

for both differential equations above when \mu~=~0, and \mu~\in~\{1,~2,~3\}.

Lawrence B. Crowell · Oct 17, 2008

The S-matrix is a topological approach to quantum field theory. It is the jumping off point for the Veneziano amplitude and the basis for string theory. It is based on the notion of an order for the relationship between different quantum fields, which in a more standard QFT approach is a time ordered product. The physical S-matrix obeys unitarity , and in the case of an ordering the vertices are defined so they do not exchange position “freely.” Any change in the ordering of vertices are done according to some discrete symmetry, such as transitions between S-T-U amplitudes.

The ordered S-matrix is constructed so that each vertex, or particle, has a neighbor. In a linear chain for instance a general state is an S-matrix channel of the form

 |\phi\rangle=~|p_1,~\dots,~p_i,~\dots,~p_j,~\dots,~p_n\rangle 

This state or S-matrix channel is related to but distinction from the channel

 |\phi\rangle=~|p_1,~\dots,~p_j,~\dots,~p_i,~\dots,~p_n\rangle 

The best way to see this is that the particles or vertices p_i and p_j have exchanged their neighbors, and a certain "relationship” structure to the amplitude has been fundamentally changed. The S-matrix is written according to S~=~1~-~2\pi T, so that given two states or channels |p_1,~\dots,~p_n\rangle and |q_1,~\dots,~q_n\rangle will be related to each other by the S-matrix as

 \langle p_1,~\dots,~p_n|S|q_1,~\dots,~q_n\rangle~=~\langle p_1,~\dots,~p_n|(1~-~2\pi T)|q_1,~\dots,~q_n\rangle 
 ~=~\langle p_1,~\dots,~p_n|q_1,~\dots,~q_n\rangle~-~2\pi\langle p_1,~\dots,~p_n|T|q_1,~\dots,~q_n\rangle. 

For the \langle-| as the in channel and |-\rangle as the out channel then it is clear that p_n and q_1 are neighbors, plus as neighbors through the T-matrix. This then eliminates an open vertex in the chain. The vertices or particles p_1 and q_n are the open elements in the chain and define the “anchor” for the chain, and are thus defined as neighbors in this manner. Hence this process defines a complete linear chain, which is similar in its structure to a gauge-“Moose,” which is a cycle of gauge fields on a compactified space, such as a Calabi-Yau space.

Each of these elements p_i defines a particle or vertex completely according to some set of quantum numbers. Thus each p_i is defined by a vector space V, which is physically some Hilbert space. The linear chain here is then an ordering on a total Hilbert space {\cal H}~=~\otimes_i V_i. Since this construction is based upon the relationship between a p_i and p_{i+1} there is then some natural bilinear operation of the form

 [-,~-]:V\times V~\rightarrow~V. 

This defines some product structure for the change in position of any of these elements. In order for this bilinear operation to describe physical states it must obey the Jacobi identity. This then requires that the vector space by k-equipped so that the bilinear operation is defined according to an isomorphism on the vector space {\cal H}~=~k\times V, where the modulus |k| is the number of elements in the chain. We then have that the isomorphism

 Y:{\cal H}\times{\cal H}~\rightarrow~{\cal H}\times{\cal H} 
 Y\big((x,~p)\otimes(y,~q)\big)~=~(x,~p)\otimes(y,~q)~+~(1,~0)\otimes(0,~[p,~q]). 

The application of Y\otimes id on {\cal H}\times{\cal H}\times{\cal H} then gives

 Y\otimes id\big((x,~p)\otimes(y,~q)\otimes(z,~r)\big)~=~(x,~p)\otimes(y,~q)\otimes(z,~r)~+~ 
 (1,~0)\otimes(0,~[[p,~q],~r]~+~[[q,~r],~p]~+~[[r,~p],~q]). 

This isomorphism on the three spaces is the Yang-Baxter equation, which is satisfied if the permuted double commutator sum vanishes, which is the Jacobi equation. The elements p,~q,~r could be the momentum operators D~+~iA, and the Jacob identity the conservation law

 [[D_a,~D_b],~D_c]~=~\epsilon_{abcd}D_eF^{de}~=~0. 

The Yang-Baxter equation is satisfied by the following commutative diagram:

The Yang-Baxter relationship is then defined in the S-matrix by the following observation. The optical theorem with S~=~1~-~2\pi T and the projection of the density matrix according to

 \rho^\prime~=~S\rho S^\dagger~=~\rho~+~2\pi i[T,~\rho]. 

The neighborhood rule tells us that the commutator is then between elements of the |-\rangle and the \langle-| with regards to the transition or T-matrix, and is thus an example of a neighbor exchange rule.

The Yang-Baxter equation describes braids, which are in this case links between nodes, vertices or particles in the chain. These links are the exchanged with each other just as a braid can connect links at different points in an array. We consider the four nodes with the braids indicated in figure 2. The composition of these braids then leads to the following rules. Every braid in the braid group B_4 is a composition of a number of these braids and their inverses. These three braids are generator of the group B_4. An braid is read from left to right; whenever a crossing of strands i and i + 1 occurs, s_i or s_i^{-1} is written down, if strand i moves under or over strand i + 1 respectively. Upon reaching the right hand end, the braid has been written as a product of the σ's and their inverses.
It is clear that

 s_1s_3~=~s_3s_1 

in addition there are the relations which are not quite as obvious:

 s_1s_2s_1~=~s_2s_1s_2,~s_2s_3s_2~=~s_3s_2s_3 

These last two are specific examples of the Yang-Baxter equation for braids, which is in general s_is_{i+1}s_i~=~s_{i+1}s_is_{i+1}.

{\bf Associahedra and Configuration Spaces of the S-matrix}

The Yang-Baxter equation describes braids, which are compositions of paths. In general the theory here needs to be extended to compositions of loops. The S-matrix acts upon a loop composed of \langle-| and |-\rangle to define the composition of two loops \langle-|-\rangle with 2\pi i\langle-|T|-\rangle. Homotopy is the mathematical theory for loop topology. For a topological space (X,~p) the loop space \Omega X is defined by the continuous map

 \phi:[0,~1]~\rightarrow~X, 

with the compact open-set topology on the endpoints \phi(0)~=~\phi(1)~=~p. Here the vertex or particle p is considered to be the base point of the map. The composition or multiplication of points is given by the rule,

 \pi_1:\Omega X\times \Omega X~\rightarrow~\Omega X, 

as the composition of two such maps parameterized by t~\in~[0,~1] by

 \pi_1(\phi,~\psi)(t)~=~\phi\cdot\psi(t)~=~\left\{\matrix{\phi(2t),& 0,~\le~t~\le~1/2\cr\psi(2t~-~1),& 1 /2~\le~t~\le~1}\right\} 

The interval [0,~1] is the interval which maps two loops and then labelled as K^2. Similarly the composition of three loops may be defined a second homotopy group according to the map between two compositions of three maps

 \pi_2(\phi\cdot\psi)\cdot\chi~\rightarrow~\phi\cdot(\psi\cdot\chi), 

given by the map

 \pi_2:K^3\times(\Omega X)^3~\rightarrow~\Omega X, 

with K^2~=~[0,~1]^2

 \pi_2(0,~-,~-,~-)~=~\pi_1(\pi_1(-,~-),~-),~\pi_2(1,~-,~-,~-)~=~\pi_1(-,~\pi_1(-,~-)) 

This map or composition may be written for two parameters t,~u as

 \pi_2(\phi,~\psi,~\chi)(t,~u)~=~ 
 \left\{\matrix{\phi((4~-~2u)t),& 0,~\le~t~\le~1/(1~-~2u)\cr\psi(4t~-~(1~+~u)),& 1 /(4~-~2u)~\le~t~\le~1/4~+~1/(4~-~2u)\cr \chi((1~+~u)(2t~-~(4~-~u)/(4~-~2u)),& 1/ 4~+~1/(4~-~2u)~\le~t~1}\right\} 

Of course this can be extended further! In this case the parameter space is K^4 and \partial K^4, for K^4 a two dimensional space. This defines the 3-homotopy or compositions of two dimensional surfaces. Homotopy parameterized by \partial K^3 is constructed from compositions of \pi_1 and \pi_3, while \pi_2 parameterized by K^4 is constructed from four continuous maps. These are a set of maps

 \pi_3:K^4\times(\Omega X)^4~\rightarrow\Omega X,~\otimes_i\prod_i\partial K^4~\rightarrow~\Omega X 

where the second map is composition of the lower homotopy groups on the boundary of the pentagon K^4, and the product is a sum over the indices on \pi_i which equal 3. This results in the homotopy groups associated with the pentagon K^4 and its boundary \partial K^4~\simeq~S^1 as seen in figure 3

This of course may be extended to arbitrary homotopy groups according to

 \pi_{n-1}:K^n\times(\Omega X)^n~\rightarrow~\Omega X, 

which defines a polytope K^5. This defines the vertices according to compositions of \pi_1, lines according to compositions of \pi_2 with two \pi_1 groups, faces according to compositions of \pi_3 with a \pi_1, or a \pi_2 and a \pi_2, and the volumes as given by \pi_4. The homotopy elements are computed, or found, rather tediously. The vertices of the polytope are

 \pi_1(\pi_1(\pi_1(\pi_1(-,-),-),-),-),~\pi_1(\pi_1(\pi_1(-,\pi_1(-,-)),-),-), 
 \pi_1(\pi_1(-,\pi_1(\pi_1(-,-),-)),-),~\pi_1(-,\pi_1(\pi_1(\pi_1(-,-),-),-)), 
 \pi_1(\pi_1(-,\pi_1(-,\pi_1(-,-))),-),~\pi_1(-,\pi_1(-,\pi_1(\pi_1(-,-),-))), 
 \pi_1(-,\pi_1(\pi_1(-\pi_1(-,-)),-)),~\pi_1(-,\pi_1(-,\pi_1(-\pi_1(-,-)))), 
 \pi_1(\pi_1(\pi_1(-,-),(\pi_1(-,-)),-),~\pi_1(\pi_1(-,-),(\pi_1(\pi_1(-,-),-))) 
 \pi_1(\pi_1(-,-),\pi_1(-,\pi_1(-,-))),~\pi_1(\pi_1(-,-),(-,\pi_1(\pi_1(-,-))) 
 \pi_1(-,\pi_1(-,\pi_1(\pi_1(-,-),-)),~\pi_1(-,\pi_1(\pi_1(-,-),\pi_1(-,-))) 

The edgelinks of the polytope are similarly

 \pi_2(\pi_1(\pi_1(-,-),-),-,-),~\pi_2(\pi_1(-,\pi_1(-,-)),-,-),~\pi_2(-,\pi_1(\pi_1(-,-),-),-), 
 \pi_2(-,\pi_1(-,\pi_1(-,-),-),~\pi_2(\pi_1(-,-),\pi_1(-,-),-),~\pi_2(-,\pi_1(-,-),\pi_1(-,-)), 
 \pi_2(\pi_1(-,-),(-,\pi_1(-,-))),~\pi_2(-,\pi_1(-,\pi_1(-,-),-),~ \pi_2(-,-,\pi_1(\pi_1(-,-),-)), 
 \pi_2(-,-,\pi_1(-,\pi_1(-,-))),~\pi_1(-,\pi_1(-,\pi_2(-,-,-)),~\pi_1(\pi_2(-,-,-),\pi_1(-,-)), 
 \pi_1(\pi_1(-,-),\pi_2(-,-,-)),~\pi_1(\pi_2(-,-,-),\pi_1(-,-)),~~\pi_1(\pi_1(-,\pi_2(-,-,-)),-) 
 \pi_1(\pi_2(\pi_1(-,-),-,-),-),~\pi_1(\pi_2(-,\pi_1(-,-),-),-),~\pi_1(\pi_2(-,-,\pi_1(-,-)),-) 
 \pi_1(-,\pi_2(-,-,\pi_1(-,-))),~\pi_1(-,\pi_2(-,\pi_1(-,-),-))),~\pi_1(-,\pi_2(\pi_1(-,-),-,-)) 

Finally the terms corresponding to the faces or plaquettes of the polytope are

 \pi_3(\pi_1(-,-),-,-,-),~\pi_3(-,\pi_1(-,-),-,-),~\pi_3(-,-,\pi_1(-,-),-) 
 \pi_3(-,-,-,\pi_1(-,-)),~\pi_1(\pi_3(-,-,-,-),-),~\pi_1(-,\pi_3(-,-,-,-)) 
 \pi_2(\pi_2(-,-,-),-,-),~\pi_2(-,\pi_2(-,-,-),-),~\pi_2(-,-,\pi_2(-,-,-)) 

Finally the body of the polytope is given by \pi_4(-,-,-,-,-) There are 14 vertices, 21 edgelinks and 9 plaquettes, which gives the Euler characteristic \chi~=~V~-~E~+~P =~14~-~21~+~9 =~2, which is the topology of a sphere and a proper polytope. The plaquettes of the polytope indicate the structure of polytope, where there are 6 faces composed of \pi_3 and \pi_1 and 3 composed of two \pi_2’s. This polytope is illustrated below in figure 4

This is a regular polytope, which is Stasheff’s K^5 associahedron. Each K^n is a generalization of the pentagon. A full generalization of the theory of associahedra is given by Jean-Louis Loday at:

http://www.claymath.org/programs/outreach/academy/LectureNotes05/Loday.pdf

The number of nodes in the n-1-associahedron is equivalent to the number of binary trees with n nodes, which is the Catalan number C_n.

The associahedron is the basic tool in the study of homotopy associative Hopf spaces.

Loday provides the following method for associahedron construction. Let A_n be the set of planar binary trees with n+1 leaves. Define a_n as the number of leaves to the left of the i^{th} vertex and b_i as the number of leaves to the right of the i^{th} vertex. For t~\in~A_n, define

 M(t)~=~(a_1b_1,~a_2b_2,~\dots,~a_nb_n) 

The n-1-associahedron is then defined as the convex hull of M(t).

The above construction of homotopy groups is an association system of elements. The homotopy groups are associative, but the operand they define determines associators and commutators. The pentagon determines a system of commutators and associators according to diagram 5

The associahedra then define operads as a sequence of topological spaces A_{[0]},~\dots,A_{[n]} with an elements e~\in~A_{[1]} and a multiplication map

 \mu_s:A_{[k]}\times A_{[n_1]}\times A_{[n_{k-1}]}~\rightarrow~A_{[n]} 

for the order preserving map s:[n]~\rightarrow~[k] and fibration s^{-1}(k)~\simeq~[n_l]. These obey associativity and unitary conditions on sequences of order preserving maps. This is then a generalized construction of the S-matrix according to operad compositions illustrated in figure 5

This figure gives an ordered tree-leaf configuration of order elements. The S-matrix may then be arranged according to such more general orderings and sequences determined by Catalan numbers. These combinatorial relationships will then define generalized amplitude and vertex operators such as the quark gluon interaction in figure 6 determined by the root vectors of the field theoretic group.

Lawrence B. Crowell · Oct 17, 2008

{\bf Associativity, Octonions and Skyrme Quark-like Models}

The octonions are generally nonassociative, and the associator (ab)c~-~a(bc)~=~[a,~b,~c] leads to uncertainty in the definition of Jacobi identities. However here an identification is found of the associator with a new Chern-Simons Lagrangian that determines the dynamics due to the nonassociative structure of the octonions. This further indicates that octonionic field theory has a moduli space construction with connections to the geometry of knots. The mathematics of this moduli space are only introduced in the most elementary fashion, and this is then presented as a new frontier of mathematical research.

We consider the nonassociative bosonic fields \phi_\alpha,~\phi_\beta,~\phi_\gamma and their commutators. The commutator [\phi_\alpha\phi_\beta,~ \phi_\gamma] is

 [\phi_\alpha\phi_\beta,~ \phi_\gamma]~=~(\phi_\alpha\phi_\beta)\phi_\gamma~-~\phi_\gamma(\phi_\alpha\phi_\beta) 
 ~=~(\phi_\alpha\phi_\beta)\phi_\gamma~-~(\phi_\gamma\phi_\alpha)\phi_\beta~+~[\phi_\gamma,~ \phi_\alpha,~ \phi_\beta] 
 ~=~(\phi_\alpha\phi_\beta)\phi_\gamma~-~(\phi_\alpha\phi_\gamma)\phi_\beta~-~[\phi_\gamma,~ \phi_\alpha]\phi_\beta~+~[\phi_\gamma,~ \phi_\alpha,~ \phi_\beta] 
 ~=~\phi_\alpha(\phi_\beta\phi_\gamma)~-~(\phi_\alpha\phi_\gamma)\phi_\beta~+~[\phi_\alpha,~ \phi_\beta,~ \phi_\gamma]~-~[\phi_\gamma,~ \phi_\alpha]\phi_\beta~+~[\phi_\gamma,~ \phi_\alpha,~ \phi_\beta] 
 ~=~\phi_\alpha(\phi_\gamma\phi_\beta)~-~(\phi_\alpha\phi_\gamma)\phi_\beta~+~\phi_\alpha[\phi_\beta,~ \phi_\gamma]~+~ 
 [\phi_\alpha,~ \phi_\beta,~ \phi_\gamma]~-~[\phi_\gamma,~ \phi_\alpha]\phi_\beta~+~[\phi_\gamma,~ \phi_\alpha,~ \phi_\beta] 

and so

 ~=~\phi_\alpha[\phi_\beta,~\phi_\gamma]~+~[\phi_\alpha,~\phi_\gamma]\phi_\beta~-~[\phi_\alpha,~ \phi_\beta,~ \phi_\gamma]~+~ [\phi_\gamma,~\phi_\alpha,~ \phi_\beta]~+~ [\phi_\alpha,~ \phi_\gamma,~ \phi_\beta] 

The double commutator is

 [\phi_\alpha,~[\phi_\beta,~\phi_\gamma]]~=~2\phi_\alpha[\phi_\beta,~\phi_\gamma]~+~[\phi_\alpha,~\phi_\beta]\phi_\gamma~-~[\phi_\alpha,~\phi_\gamma]\phi_\beta~-~ 
 [\phi_\alpha,~ \phi_{[\beta},~ \phi_{\gamma]}]~+~ [\phi_{[\gamma},~\phi_\alpha,~ \phi_{\beta]}~+~ [\phi_\alpha,~ \phi_{[\gamma},~ \phi_{\beta]}], 

and that the Jacobi identity is of the form

 [\phi_\alpha,~[\phi_\beta,~\phi_\gamma]]~+~ [\phi_\beta,~[\phi_\gamma,~\phi_\alpha]]~+~ [\phi_\gamma,~[\phi_\alpha,~\phi_\beta]]~=~ 
 [\phi_\alpha,~\phi_{[\beta},~\phi_{\gamma]}]~+~[\phi_\beta,~\phi_{[\gamma},~\phi_{\alpha]}]~+~[\phi_\gamma,~\phi_{[\alpha},~\phi_{\beta]}], 

where the cyclicity of the associator has been used. Here the subscripts [ and ] indicate anticommutation of these indices, and the associator (~) are defined by the commutator.

If the fields \phi_{[\alpha}\phi_{\beta]}~=~F_{\alpha\beta}, where these field \phi_\alpha are thought of as the differential operator {\cal D}_\alpha and the F_{\alpha\beta} are components of the field strength tensor, then the above Jacobi identity is reduced to

 Cyc[\phi_\alpha,~[\phi_\beta,~\phi_\gamma]]~=~\phi_\alpha F_{\beta\gamma}~+~\phi_{[\beta}F_{\gamma]}\alpha~+~\phi_{[\gamma}F_{\alpha\beta]}~-~ 
 (\phi_\alpha\phi_\beta)\phi_\gamma~-~(\phi_\beta\phi_\gamma)\phi_\alpha~-~(\phi_\gamma\phi_\alpha)\phi_\beta 

The nonassociativity of octonions means this involves a \lq\lq torsion\rq\rq. Using permutation symmetry of the associator an alternate Jacobi identity is

 Cyc[[\phi_\alpha,~\phi_\beta],~\phi_\gamma]~=~ F_{\alpha\beta}\phi_\gamma~+~ F_{\beta\gamma}\phi_\alpha~+~ F_{\gamma\alpha}\phi_\beta ~-~ 
 \phi_{[\alpha}(\phi_\beta]\phi_\gamma)~-~\phi_{[\beta}(\phi_{\gamma]}\phi_\alpha)~-~\phi_{[\gamma}(\phi_{\alpha]}\phi_\beta). 

The Jacobi identity is then imposed as a topological requirement that d^2~=~0, where the fields \phi_\alpha are coboundary operators. With the antisymmetry of the associator we have the difference between these two permuting commutators by the Jacobi identity requires that

 Cyc[\phi_\alpha,~[\phi_\beta,~\phi_\gamma]]~+~ Cyc[\gamma,~[\phi_\alpha,~\phi_\beta]]~+~ Cyc[[\phi_\gamma,~\phi_\alpha],~\phi_\beta]~=~0, 

which means that the associator may be identified as

 [\phi_{[\alpha},~\phi_{\beta]},~\phi_\gamma]~=~{1\over 2}[\phi_\alpha,~ F_{\beta\gamma}] 

This indicates that the associator [\phi_\alpha,~\phi_\beta,~\phi_\gamma] is then a determinant of a part of the Chern-Simons Lagrangian

 {\cal L}_{cs}~=~{\cal L}^0_{cs}~+~ {\bf\Lambda}_{[\phi_\alpha,~\phi_\beta,~\phi_\gamma]}~=~\int (A\wedge dA~-~{2\over 3}A\wedge A\wedge A)~+~\int{[\phi_\alpha,~\phi_\beta,~\phi_\gamma]} 

Here {\cal L}^0_{cs} pertains to dynamics of with the seven associated subspaces, and {\bf\Lambda}_{[\phi_\alpha,~\phi_\beta,~\phi_\gamma]} involves dynamics from nonassociative action of coboundary operators. This is an interesting find that has connections to knot theory. Here the Chern-Simons Lagrangian from the associator determines the dynamics involved with products of connection terms that are nonassociative.

{\bf Knot Theory, Hopf fibration and Octonions}

Consider \cal M to be out 7 dimensional manifold. We then assign a vector space to this by Z({\cal M}). Consider the vectors X and Y where XY is the Fano plane product. There are then a number of possible properties of this vector space

 Z({\cal M}^*) ~=~ Z({\cal M})^* ~duality 
 Z(XY) ~=~ Z(X)\otimes Z(Y) 

and in the latter case if X and Y are generalized to disjoint subsets of S^7 then

 Z(S_1\vee S_2)~ =~ Z(S_1)\otimes Z(S_2), 

where "\vee" is a product analogous to the wedge product between differential forms. So let V_1, V_2, V_3 be disjoint subsets of S^7, and let the boundaries of these be

 \partial V_1~=~S_1~+~S_2 
 \partial V_2~=~S_2~+~S_3 
 \partial V_3~=~S_3~+~S_4 

Then Z(V_1) is a homomorphism Z(S_1)~\rightarrow~Z(S_2) and Z(V_2) is a homomorphism Z(S_2)~\rightarrow~Z(S_3) etc. We then have that

 Z(V_1 \cup V_2) ~=~ Z(V_1)Z(V_2) \in Hom(Z(S_1), ~Z(S_2)). 

This leads to an interesting result

 Z((V_1 \cup V_2) \cup V_3) ~=~ (Z(V_1)Z(V_2))Z(V_3) ~\in~ Hom(Hom(Z(S_1), ~Z(S_2)), ~Z(S_3))) 
and
 Z(V_1 \cup (V_2 \cup V_3)) ~=~ Z(V_1)(Z(V_2)Z(V_3))~\in~ Hom(Z(S_1), ~Hom(Z(S_2), ~Z(S_3))), 

and nonassociativity means that

 Z((V_1 \cup V_2) \cup V_3) ~\ne~ Z(V_1 \cup (V_2 \cup V_3)) 
and
 Hom(Hom(Z(S_1), ~Z(S_2)), ~Z(S_3)))~\ne~Hom(Z(S_1), ~Hom(Z(S_2),~ Z(S_3))) 

This leads to some interesting generalizations of the constructions used in topology. This generalizes the traditional atlas-chart construction with transition functions. This construction is a bit more complex than usual, for one has to consider the overlap of 3 charts to get associative and nonassociative parts. There will then be transition functions for elements. We have the duality

 \langle e_i,~ \omega_j\rangle ~=~ \delta_{ij} 

and differential forms are covariant d~\rightarrow~d ~+~ iA. The space of connections and the moment map \{A\}~\rightarrow~ F for gauge fields is the moduli space A/{\cal G}, \cal G = Lie group.

These correspond to the set of gauge connections that are equivalent by a gauge choice. The two-form for the curvature F ~=~ dA is then the set of fields. We have dF ~=~ 0, and so F is closed and this defines a conservation law. Let a be a 1-form on A/{\cal G}, so that B ~=~ da and dB ~=~ 0. One could have a~ =~ 0 for a simply connected region. However, for multiply connected A/{\cal G} and may be a closed/exact form. This is the Chern-Simons form. For an abelian theory this leads to a ~=~ A\wedge dA,

 L(A) ~= \int A \wedge dA, 
and for the nonabelian theory
 L(A) ~=~ \int(A \wedge dA ~+~ {2\over 3}A \wedge A \wedge A) = \int A \wedge DA 

The integrand is the Chern-Simons form and the integral is the Hopf invariant. This integral then enters into the path integral

 Z[A] ~=~ \int {\cal D}[A] exp(iL(A)). 

{\bf Discussion}

We have illustrated a connection between the S-matrix, associahedra and knot theory. The underlying nonassociativity of quantum states, and their corresponding associative homotopy groups, illustrates that a more general method for ordering states is possible with an underlying nonassociativity. The result is a field theory which is similar to a Skyrme model. This also can produce a generalized system for computing vertex operators. The Chern-Simon’s Lagrangian results in actions for coherent states, which under a product of all vacuum configurations gives a beta function for a four-point vertex function. In a similar manner the four point vertex function corresponds to ordered arrangements of four nodes. All possible combinations of these elements is a binomial distribution. However, the tree leaf ordering on these combinations according to a system of preferential attachments. This will define an Euler Beta function, which is the Veneziano amplitude result.

\bye

Information Preservation in Quantum Gravity

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