Open access peer-reviewed chapter - ONLINE FIRST

Stueckelberg-Horwitz-Piron Canonical Quantum Theory in General Relativity and Bekenstein-Sanders Gauge Fields for TeVeS

By Lawrence P. Horwitz

Submitted: February 12th 2019Reviewed: June 19th 2019Published: July 18th 2019

DOI: 10.5772/intechopen.88154

Downloaded: 50

Abstract

A consistent (off-shell) canonical classical and quantum dynamics in the framework of special relativity was formulated by Stueckelberg in 1941 and generalized to many-body theory by Horwitz and Piron in 1973 (SHP). This theory has been embedded into the framework of general relativity (GR), here denoted by SHPGR. The canonical Poisson brackets of the SHP theory remain valid (invariant under local coordinate transformations) on the manifold of GR and provide the basis for formulating a canonical quantum theory. The relation between representations based on coordinates and momenta is given by Fourier transform; a proof is given here for this functional relation on a manifold. The potential which may occur in the SHP theory emerges as a spacetime scalar mass distribution in GR. Gauge fields, both Abelian and non-Abelian on the quantum mechanical SHPGR Hilbert space in both the single particle and many-body theory, may be generated by phase transformations. Application to the construction of Bekenstein and Sanders in their solution to the lensing problem in TeVeS is discussed.

Keywords

  • relativistic dynamics
  • general relativity
  • quantum theory on curved space
  • non-Abelian gauge fields
  • Bekenstein-Sanders field
  • TeVeS

1. Introduction

The relativistic canonical Hamiltonian dynamics of Stueckelberg, Horwitz, and Piron (SHP) [1] with scalar potential and gauge field interactions for single- and many-body theories can, by local coordinate transformation, be embedded into the framework of general relativity (GR). This embedding provides a basis for the work of Horwitz et al. [2, 3] in their discussion of the origin of the field introduced by Bekenstein and Sanders [4] to explain gravitational lensing in the TeVeS formulation of modified Newtonian dynamic (MOND) theories [5, 6, 7, 8, 9, 10].

The theory was originally formulated for a single particle by Stueckelberg in [11, 12, 13]. Stueckelberg envisaged the motion of a particle along a world line in spacetime that can curve and turn to flow backward in time, resulting in the phenomenon of pair annihilation in classical dynamics. The world line was then described by an invariant monotonic parameter τ. The theory was generalized by Horwitz and Piron in [14] (see also [15, 16]) to be applicable to many-body systems by assuming that the parameter τis universal (as for Newtonian time, enabling them to solve the two-body problem classically, and later, a quantum solution was found by Arshansky and Horwitz [17, 18, 19], both for bound states and scattering theory).

Performing a coordinate transformation to general coordinates, along with the corresponding transformation of the momenta (the cotangent space of the original Minkowski manifold), one obtains [20] the SHP theory in a curved space of general coordinates and momenta with a canonical Hamilton-Lagrange (symplectic) structure. We shall refer to this generalization as SHPGR. We discuss the extension of the Abelian gauge theory described in Ref. [20] to the non-Abelian gauge discussed in [2, 3].

The invariance of the Poisson bracket under local coordinate transformations provides a basis for the canonical quantization of the theory, for which the evolution under τis determined by the covariant form of the Stueckelberg-Schrödinger equation [1].

In this chapter, we assume a τ-independent background gravitational field; the local coordinate transformations from the flat Minkowski space to the curved space are taken to be independent of τ, consistently with an energy momentum tensor that is τindependent. In a more dynamical setting, when the energy momentum tensor depends on τ, the spacetime is evolved nontrivially [20, 21].

2. Embedding of single particle dynamics with external potential in GR

We write the SHP Hamiltonian [1, 11, 12, 13] as

K=12Mημνπμπν+VξE1

where ημνis the flat Minkowski metric +++and πμ,ξμare the spacetime canonical momenta and coordinates in the local tangent space of a general manifold, following Einstein’s use of the equivalence principle.

The existence of a potential term (which we assume to be a Lorentz scalar), representing nongravitational forces, implies that the “free fall” condition is replaced by a local dynamics carried along by the free falling system (an additional force acting on the particle within the “elevator” according to the coordinates in the tangent space).

The canonical equations are

ξ̇μ=Kπμπ̇μ=Kξμ=Vξμ,E2

where the dot here indicates d, with τthe invariant universal “world time.” Since then

ξ̇μ=1Mημνπν,orπν=ηνμMξ̇μ,E3

the Hamiltonian can then be written as

K=M2ημνξ̇μξ̇ν+Vξ.E4

It is important to note that, as clear from (3), that ξ̇0=dthas a sign opposite to π0which lies in the cotangent space of the manifold, as we shall see in the Poisson bracket relations below. The energy of the particle for a normal timelike particle should be positive (negative energy would correspond to an antiparticle [1, 11, 12, 13]). The physical momenta and energy therefore correspond to the mapping

πμ=ημνπμ,E5

back to the tangent space. Thus, equivalently, from (2), ξ̇μ=1/Mπμ. This simple observation will be important in the discussion below of the dynamics of a particle in the framework of general relativity, for which the metric tensor is nontrivial.

We now transform the local coordinates (contravariantly) according to the diffeomorphism

dξμ=ξμxλdxλE6

to attach small changes in ξto the corresponding small changes in the coordinates xon the curved space, so that

ξ̇μ=ξμxλẋλ.E7

The Hamiltonian then becomes

K=M2gμνẋμẋν+Vx,E8

where Vxis the potential at the point ξcorresponding to the point x(a function of ξin a small neighborhood of the point x) and

gμν=ηλσξλxμξσxνE9

Since Vhas significance as the source of a force in the local frame only through its derivatives, we can make this pointwise correspondence with a globally defined scalar function Vx.1

The corresponding Lagrangian is then

L=M2gμνẋμẋνVx,E10

In the locally flat coordinates in the neighborhood of xμ, the symplectic structure of Hamiltonian mechanics implies that the momentum2πμ, lying in the cotangent space of the manifold ξμ, transforms covariantly under the local transformation (5), that is, as does ξμ, so that we may define

pμ=ξλxμπλ.E11

This definition is consistent with the transformation properties of the momentum defined by the Lagrangian (10):

pμ=Lxẋẋμ,E12

yielding

pμ=Mgμνẋν.E13

The second factor in the definition (9) of gμνin (13) acts on ẋν; with (7) we then have (as in (11))

pμ=Mηλσξλxμξ̇σ=ξλxμπλ.E14

As we have remarked above for the locally flat space in (5), the physical energy and momenta are given, according to the mapping,

pμ=gμνpν=MẋνE15

back to the tangent space of the manifold, which also follows directly from the local coordinate transformation of (3) and (5).

It is therefore evident from (15) that

ṗμ=Mx¨μ.E16

We see that ṗμ, which should be interpreted as the force acting on the particle, is proportional to the acceleration along the orbit of motion (a covariant derivative plus a gradient of the potential), as described by the geodesic-type relation. This Newtonian-type relation in the general manifold reduces in the limit of a flat Minkowski space to the corresponding SHP dynamics and in the nonrelativistic limit, to the classical Newton law. We remark that this result does not require taking a post-Newtonian limit, the usual method of obtaining Newton’s law from GR.

We now discuss the geodesic equation obtained by studying the condition

ξ¨μ=1Mπ̇μ=1MημνVξξν.E17

To do this, we compute

ξ¨μ=dξμxλẋλ=2ξμxλxγẋγẋλ+ξμxλx¨λ=1MημνxλξνVxxλ,E18

so that, after multiplying by xσξμand summing over μ, we obtain

x¨σ=xσξμ2ξμxλxγẋγẋλ1MημνxλξνxσξμVxxλ.E19

Finally, with (9) and the usual definition of the connection

Γσλγ=xσξμ2ξμxλxγE20

we obtain the modified geodesic-type equation

x¨σ=Γσλγẋγẋλ1MgσλVxxλ,E21

from which we see that the derivative of the potential Vξis mapped, under this coordinate transformation into a force resulting in a modification of the acceleration along the geodesic-like curves, that is, (16) now reads

ṗμ=Mx¨ν=MΓσλγẋγẋλgσλVxxλE22

The procedure that we have carried out here provides a canonical dynamical structure for motion in the curvilinear coordinates. We now remark that the Poisson bracket remains valid for the coordinates xp. In the local coordinates ξπ, the τderivative of a function Fξπis

dFξπ=Fξπξμξ̇μ+Fξππνπ̇μ=FξπξμKπμFξππμKξνFKPBξπ.E23

If we replace in this formula

ξμ=xλξμxλπμ=ξμxλpλ,E24

we immediately (as assured by the invariance of the Poisson bracket under local coordinate transformations) obtain

dFξπ=FxμKpμFpμKxνFKPBxpE25

In this definition of Poisson bracket, we have, as for the ξμ,πνrelation,

xμpνPBxp=δμν.E26

The Poisson bracket of xμwith the (physical energy momentum) tangent space variable pμhas then the tensor form

xμpνPBxp=gμν.E27

In the flat space limit, this relation reduces to the SHP bracket,

ξμπνPBξπ=ημν.E28

Continuing our analysis with pμ(we drop the xplabel on the Poisson bracket henceforth),

pμFxPB=Fxμ,E29

so that pμacts infinitesimally as the generator of translation along the coordinate curves and

xμFpPB=Fppμ,E30

so that xμis the generator of translations in pμ. In the classical case, if Fpis a general function of pμ, we can write at some point x,3

xμFpPB=gμνxFppν.E31

This structure clearly provides a phase space which could serve as the basis for the construction of a canonical quantum theory on the curved spacetime.

We now turn to a discussion of the dynamics introduced into the curved space by the procedure outlined above.

We may also write (22) in terms of the full connection form by noting that with (9),

gλγxμ=ηαβ2ξαxλxμξβxγ+ξαxλ2ξβxγxμ.E32

Multiplying by ẋγẋλ, the two terms combine to give a factor of two. We then return to the original definition of Γin (20) in the form

2ξαxλxμ=ξαxσΓσλμ,E33

so we can write

gλγxμẋγẋλ=2ηαβξαxσξβxγΓσλμẋγẋλ=2gσγΓσλμẋγẋλ.E34

We therefore have

ṗμ=Vxxμ+MgσγΓσλμẋγẋλ.E35

3. Quantum theory on the curved space

The Poisson bracket formulas (25) and (26) can be considered as a basis for defining a quantum theory with canonical commutation relations

xμpν=iδμν,E36

so that

pμFx=iFxμ,E37

and

xμFp=iFppμ.E38

The transcription of the Stueckelberg-Schrödinger equation for a wave function ψτxcan be taken to be (see also [26, 27, 28])

iτψτx=Kψτx,E39

where the operator valued Hamiltonian can be taken to be the Hermitian form (42), written below, on a Hilbert space defined with scalar product (with invariant measure; we write g=detgμν),

ψχ=d4xgψτxχτx.E40

To construct a Hermitian Hamiltonian, we first study the properties of the canonical momentum in coordinate representation. Clearly, in coordinate representation, ixμis not Hermitian due to the presence of the factor gin the integrand of the scalar product. The problem is somewhat analogous to that of Newton and Wigner [29] in their treatment of the Klein-Gordon equation in momentum space. It is easily seen that the operator

pμ=ixμi21gxxμgxE41

is essentially self-adjoint in the scalar product (40), satisfying as well as the commutation relations (36).4

Since pμis Hermitian in the scalar product (41), we can write the Hermitian Hamiltonian as

K=12Mpμgμνpν+Vx,E42

consistent with the local coordinate transformation of (1). The integration (40) must be considered as a total volume sum with invariant measure on the whole space, consistent with the notion of Lebesgue measure and the idea that the norm is the sum of probability measures on every subset contained. We return to this point in our discussion of the Fourier transform below.

4. Canonical quantum theory and the Fourier transform

To complete the construction of a canonical quantum theory on the curved space of GR, we discuss first the formulation of the Fourier transform fxf˜pfor a scalar function fx(we shall use xμand the canonically conjugate pμin this discussion). Let us define (gdetgμν)

f˜p=d4xgxeipμxμfx.E43

The inverse is given by

eipμxμf˜pd4p=d4peipμxμxμfxgxd4x=2π4fxgxE44

so that

f˜p=12π4gxeipμxμf˜pd4p.E45

One sees immediately that under diffeomorphisms, for which with the scalar property fx=fx, f˜pf˜p. The Fourier transform of fxis

f˜p=d4xgxeipμxμfx,E46

By change of integration variables, we have

f˜p=d4xgxeipμxμfx,E47

In Dirac notation,

fx=<xf>,E48

and we write as well

f˜p=<pf>.E49

For

<xp>=12π4gxeipμxμ<px>=gxeipμxμ,E50

we have, for example, the usual action of transformation functions

<xp><pf>d4p=<xf>,E51

where we have used

<xp><px>d4p=12π4gxd4peipμxμeipμxμgx=δ4xx.E52

Note that the transformation functions <xp>and <px>are not simple complex conjugates of each other, but require nontrivial factors of gxand its inverse to satisfy the necessary transformation laws on the manifold. Conversely (the factors gxand its inverse cancel), we should obtain

<px><xp>d4x=δ4pp.E53

The validity of (53) is not obvious on a curved space. We therefore provide a simple, but not trivial, proof of (53). For

d4peipμxμxμ=2π4δ4xxgE54

we must have

f˜p=12π4d4xd4peipμpμxmuf˜p,E55

that is, exchanging the order of integrations, on the set f˜p,

Δpp=12π4d4xeipμpμxμ=δ4pp.E56

We now represent the integral as a sum over small boxes around the set of points xBthat cover the space and eventually take the limit as for a Riemann integral.5 In each small box, the coordinatization arises from an invertible transformation from the local tangent space in that neighborhood. We write

xμ=xBμ+ημboxBE57

where

ημ=xμξλξλE58

and ξλis in the flat local tangent space at xB.

We now write the integral (56) as

Δpp=12π4ΣBBd4ηeipμpμxBμ+ημ=12π4ΣBeipμpμxBμBd4ηeipμpμημ.E59

Let us call

IB=Bd4ηeipμpμημ.E60

In this neighborhood, call

xμξλ=aμλB,E61

which we assume a constant matrix (Lorentz transformation) in each box. In (60), we then have

IB=Bd4ξdetaeipμpμaμλBξλ.E62

However, we can make a change of variables; we are left with

IB=Bd4ξeipμpμξμ.E63

in each box.

However, the transformations aμλBin the neighborhood of each point Bmay be different, and therefore the set of transformed boxes may not cover (boundary deficits) the full domain of spacetime coordinates (one can easily estimate that the deficit from an arbitrarily selected set can be infinite in the limit).

We may avoid this problem by assuming geodesic completeness of the manifold and taking the covering set of boxes, constructed of parallel transported edges, along geodesic curves. Parallel transport of the tangent space boxes then fills the space in the neighborhood of the geodesic curve we are following, and each infinitesimal box may carry an invariant volume (Liouville-type flow) transported along a geodesic curve. For successive boxes along the geodesic curve, since the boundaries are determined by parallel transport (rectilinear shift in the succession of local tangent spaces), there is no volume deficit between adjacent boxes.

We may furthermore translate a geodesic curve to an adjacent geodesic by the mechanism discussed in [32], so that boxes associated with adjacent geodesics are also related by parallel transport. In this way, we may fill the entire geodesically accessible spacetime volume.

Let us assign a measure to each point B:

ΔμBppIB.E64

We may then write (59) as

Δpp=12π4ΣBeipμpμxBμΔμBpp,E65

Our construction has so far been based on elements constructed in the tangent space in the neighborhood of each point B. Relating all points along a geodesic to the corresponding tangent spaces and putting each patch in correspondence by continuity, we may consider the set xBto be in correspondence with an extended flat space ξ, for which xBξBto obtain6

Δpp=12π4ΣBeipμpμξBμΔμξBpp,E66

In the limit of vanishing spacetime box volume, this approaches the Lebesgue-type integral on a flat space:

Δpp=12π4eipμpμξμξpp.E67

If the measure is differentiable, we could write

ξpp=mξppd4ξ.E68

Since the kernel Δppis to act on elements of a Hilbert space f˜p, the support for pvanishes, so that ppis essentially bounded, as we discuss below. In the small box, say, size ϵ,

ϵ/2ϵ/2dξ0dξ1dξ2dξ3eipμpμξμ=2i4Πj=0j=3sinpjpjϵ2pjpjϵ4d4ξ,E69

so that mξpp=1, and we have

Δpp=12π4eipμpμξμd4ξ,E70

or7

Δpp=δ4pp.E71

It is clear that the assertion (69) requires some discussion. For ϵ0we must be sure that pdoes not become too large, so that our local measure is equivalent to d4ξ. In one of the dimensions, what we want to find are conditions for which

sinpϵpϵE72

for ϵ0, where we have written pfor pp. As a distribution, on functions gp, the left member of (72) acts as

Gϵsinpϵpgp.E73

The function Gϵis analytic in the neighborhood of ϵ=0if pngphas a Fourier transform for all nand the series is convergent in this neighborhood, since G0is identically zero and successive derivatives correspond to the Fourier transforms of pngp(differentiating under the integral). This implies that if the (usual) Fourier transform of gpis a Cfunction (as a simple sufficient condition) in the local tangent space ξand we have appropriate convergence properties, we can reliably use the first order term in the Taylor expansion;

ddϵGϵϵ=0=cosϵpgpϵ=0E74

so that, for ϵ0,

Gϵϵg˜0,E75

where g˜ξis the Fourier transform of gp. As a distribution on such functions gp, the assertion (3.39) then follows.

5. Application to the Bekenstein-Sanders fields

We have discussed the construction of a canonical quantum theory in terms of an embedding of the SHP relativistic classical and quantum theory into general relativity. We show in this section that this systematic embedding provides a framework for the method developed by Bekenstein and Milgrom for understanding the MOND [5, 6, 7, 8, 9, 10] that appeared necessary to explain the galactic rotation curves [35].

The remarkable development of observational equipment and power of computation has resulted in the discovery that Newtonian gravitational physics leads to a prediction for the dynamics of stars in galaxies that is not consistent with observation. It was proposed that there should be a matter present which does not radiate light which would resolve this difficulty, but so far no firm evidence of the existence of such matter has emerged. Milgrom [5, 6, 7, 8, 9, 10] proposed a modification of Newton’s law (MOND) which could resolve the problem. However, since Newton’s law of gravitation emerges in the “post-Newtonian approximation” to the geodesic motion in Einstein’s theory of gravity [35], the modification of Newton’s law must involve a modification of Einstein’s theory.8 Such a modification was proposed by Bekenstein and Milgrom [5, 6, 7, 8, 9, 10] in terms of a conformal factor multiplying the usual Einstein metric.

The origin of such a conformal factor can be found in the potential term of the special relativistic SHP theory. The embedding of this theory in GR [20] brings this potential term as a world scalar. The Hamiltonian for the general relativistic case then has the form (8). It was shown by Horwitz et al. [37] that a very sensitive test by geodesic deviation can be formulated by to study stability by transforming a standard nonrelativistic Hamiltonian of the form

H=p22M+VrE76

to the form

H=12Mpigijrpj,E77

with

gijr=ϕrEEVδij,E78

that is, a conformal factor on the original metric. Applying the same idea to the Hamiltonian (8), with the gμνxof Einstein replaced by the conformal form

g˜μνx=ϕxgμνxE79

where

ϕx=kkVx,E80

with ka point in the spectrum of K, so that

H=12Mpμg˜μνxpν.E81

We see that we can in this way achieve the structure proposed by Bekenstein and Milgrom [5, 6, 7, 8, 9, 10] systematically. Moreover, in addition to providing a mechanism for achieving a realization of the MOND theory, in the original form (8), the world scalar term Vxcould represent the so-called dark energy [2, 3], establishing a relation between the MOND picture and the anomalous expansion of the universe, a question presently under study.

The theory proposed by Bekenstein and Milgrom [5, 6, 7, 8, 9, 10] did not, however, account for the lensing of light observed when light passes a galaxy. To solve this problem, Bekenstein and Sanders [4] proposed the introduction of a vector fieldnμx, satisfying the normalization constraint

nμnμ=1,E82

so that the vector is timelike.

This vector field can then be used to construct a modified meric of the form

g˜μνT=ϕgμνx+nμxnνx+ϕ1nμxnνx.E83

With this modification, Bekenstein and Sanders [4] could explain the lensing effect. In the following, we show that this new field may arise from a non-Abelian gauge transformation [38, 39] on the quantum theory that we have discussed in Section 3. Although Contaldi et al. [40] point out that a gauge field in this context can have caustic singularities due to the presence of a massive system, Horwitz et al. [2, 3] show that in the limit in which the gauge field approaches the Abelian limit, as required by Bekenstein and Sanders [4], there is a residual term that can cancel the caustic singularities.

To preserve the normalization condition (83), it is clear that we have the possibility of moving the nfield on a hyperbola with a Lorentz transformation, which we can perform by a gauge transformation.

A Lorentz transformation on nμis noncommutative, and therefore the gauge field is non-Abelian [21].

An analogy can be drawn to the usual Yang-Mills gauge on SU2, where there is a two-valued index for the wave function ψαx. The gauge transformation is a two-by-two matrix function of xand acts only on the indices α. The condition of invariant absolute square (probability) is

αβUαβψβ2=ψα2E84

Generalizing this structure, one can take the indices αto be infinite dimensional, and even continuous, so that (84) becomes (in the spectral representation for nμ)

dndnU(nn)ψ(nx)2=dnψ(nx)2,E85

implying that Unn(at each point x) is a unitary operator on a Hilbert space L2dn. Since we are assuming that nμlies on a hyperbola determined by (83), the measure is

dn=d3nn0,E86

that is, a three-dimensional Lorentz invariant integration measure.

We now examine the gauge condition:

pμϵnμ=UpμϵnμψE87

Since the Hermitian operator pμacts as a derivative under commutation relations, we obtain

nμ=UnμU1iϵUxμU1,E88

in the same form as the Yang-Mills theory [38, 39]. It is evident in the Yang-Mills theory, due to the operator nature of the second term, the field will be algebra-valued, and thus we have the usual structure of the Yang-Mills non-Abelian gauge theory. Here, if the transformation Uis a Lorentz transformation, the numerical-valued field nμwould be carried, at least in the first term, to a new value on a hyperbola. However, the second term is operator valued on L2dn, and thus, as in the Yang-Mills theory, nμwould become operator valued. Therefore, in general, the gauge field nμis operator valued.

It follows from (87) that the “field strengths”

fμν=nμxνnνxμ+iϵnμnν.E89

Under a gauge transformation nμnμ, the new fields create field strengths in the transformed form

fμν=nμxνnνxμ+iϵnμnνE90

according to

fμνx=UfμνxU1,E91

just as in the finite-dimensional Yang-Mills theories.

For

U1+iG,E92

where Gis infinitesimal, (87) becomes

nμ=nμ+iGnμ+1ϵGxμ+OG2.E93

Then,

nμnμnμnμ+inμGnμ+Gnμnμ+1ϵGxμnμ+nμGxμ.E94

Let us take

G=iϵ2ωλγnxnλnγnγnλϵ2ωλγnxNλγE95

where symmetrization is required since ωλγis a function of nas well as xand

Nλγ=inλnγnγnλ.E96

Our investigation in the following will be concerned with a study of the infinitesimal gauge neighborhood of the Abelian limit, where the components of nμdo not commute and therefore still constitute a Yang-Mills-type field. We shall show in the limit that the corresponding field equations acquire nonlinear terms and may therefore nullify the difficulty found by Contaldi et al. [40] demonstrating a dynamical instability for an Abelian vector-type TeVeS gauge field. They found that nonlinear terms associated with a non-Maxwellian-type action, such as divn2, could nullify this caustic singularity, so that the nonlinear terms we find as a residue of the Yang-Mills structure induced by our gauge transformation might achieve this effect in a natural way.

Now, the second term of (94), which is the commutator of Gwith nμnμ, vanishes, since this product is Lorentz invariant (the symmetrization in Gdoes not affect this result).

We now consider the third term in (94).

1ϵGxμnμ+nμGxμ=12ωλγxμNλγnμ+nμωλγxμNλγ=12{Nλγωλγxμnμ+ωλγxμNλγnμ+nμNλγωλγxμ+nμωλγxμNλγ}E97

There are two terms proportional to

ωλγxμnμ.

If we take

ωλγnx=ωλγkνxν,E98

where kνnν=0, then

ωλγxμnμ=kμnμωλγ=0.E99

For the remaining two terms,

nμNλγωλγxμ+ωλγxμNλγnμ=Nλγnμωλγxμ+nμNλγωλγxμ+ωλγxμnμNλγ+ωλγxμNλγnμ.E100

The commutators contain only terms linear in nμand they cancel; the remaining terms are zero, and therefore the condition nμnμ=1is invariant under this gauge transformation. It involves the coefficient ωλγwhich is a function of the projection of xμonto a hyperplane orthogonal the nμ. The vector kμof course depends on nμ. We take, for definiteness, kμ=nμnb+bμ, for some bμ0.

We now consider the derivation of field equations from a Lagrangian constructed with the ψsand fμνfμν. We take the Lagrangian to be of the form

L=Lf+Lm,E101

where

Lf=14fμνfμνE102

and

Lm=ψiτ12MpμϵnμpμϵnμΦψ+c.c.E103

In carrying out the variation of Lm, the contributions of varying the ψs with respect to nvanish due to the field equations (Stueckelberg-Schrödinger equation) obtained by varying ψ(or ψ), and therefore in the variation with respect to n, only the explicit presence of nin (103) need be taken into account.

Note that for the general case of ngenerally operator valued, we can write

ψpμϵnμpμϵnμψ=pμϵnμψpμϵnμψ,E104

since the Lagrangian density (108) contains an integration over dndn(in spectral representation, considered in lowest order, as well as an integration over dxgxin the action). In the limit in which nis evaluated in the spectral representation, and noting that pμis represented by an imaginary differential operator, we can write this as

ψpμϵnμpμϵnμψ=pμ+ϵnμ)ψpμϵnμψ,E105

that is, replacing explicitly pμby i/xμiμ(since it acts by commutator with the fields); we have

δnLm=iϵ2Mψμiϵnμψμ+iϵnμψψδnμ,E106

or

δnLm=jμnxδnμ,E107

where jμhas the usual form of a gauge invariant current.

For the calculation of the variation of Lf, we note that the commutator term in (89) is, in lowest order, a c-number function.

Calling

ωλμnλvμ,E108

we compute the variation of

nμnν=2ikνvμkμvνE109

Then, for

δnnμnν=δnγnγnμnν,E110

we compute

nγnμnν=2ikνnγvμ+kνvμnγμν.E111

With our choice of kν=nνnb+bν,

kνnγ=δνγnb+nνbγ,E112

so that

nγnμnν=2i(δνγnb+nνbγvμ+kνvμnγμν).E113

Here,

vμnγ=ωμγ+ωλμnλkσnγxσ,

so we see that

nγnμnνOγμν,E114

where the quantity Oγμνδnγdepends on the first and second derivatives of ωλμ, in general, nonlinear in nμ. We therefore have

δnnμnν=OγμνδnγE115

In the limit that ω0, its derivative and higher derivatives which appear in Oγμνmay not vanish (somewhat analogous to the case in gravitational theory when the connection form vanishes, but the curvature does not), so that this term can contribute in the limit of the an Abelian gauge.

Returning to the variation of Lf, we see that

δLf=14(μδnννδnμ+iϵδnμnμfμν+fμνμδnννδnμ+iϵδnμnμ)=νfμνδnμ+2ifμνδnμnν,E116

where we have taken into account the fact that nμnμis a c-number function and integrated by parts the derivatives of δn. We then obtain

δLf=νfμνδnμ+2iϵfλσOλσμδnμE117

Since the coefficient of δnμmust vanish, we obtain the Yang-Mills equations for the fields given the source currents:

νfμν=jμ2iϵfλσOλσμ,E118

which is nonlinear in the fields nμ, as we have seen, even in the Abelian limit, where, from (106),

jμ=iϵ2Mψμiϵnμψμ+iϵnμψψ.E119

6. Summary

In this chapter, we have shown that the formulation of MOND theory by Bekenstein and Milgrom [5, 6, 7, 8, 9, 10] can have a systematic origin within the framework of the embedding of the SHP [1] theory into general relativity [20]. The SHP formalism admits a scalar potential term that appears both in the conformal factor giving rise to the MOND functions in the galaxy and, in the original form of the Hamiltonian, to a possible candidate for “dark energy.” The solution of the lensing problem by Bekenstein and Sanders [4] by introduction of a local vector field was also shown to arise in a natural way in terms of a non-Abelian gauge field, for which, in the Abelian limit, there is a residual term that can cancel the caustic singularity found by Contaldi et al. [40] which can arise in a purely Abelian gauge theory.

Notes

  • Since Vx has the dimension of mass, one can think of this function as a scalar mass field, reflecting forces acting in the local tangent space at each point. It may play the role of “dark energy” [2, 3]. If V=0, our discussion reduces to that of the usual general relativity, but with a well-defined canonical momentum variable.
  • We shall call the quantity πμ in the cotangent space as canonical momentum, although it must be understood that its map back to the tangent space πμ corresponds to the actual physically measureable momentum.
  • In the quantized form, the factor gμνx cannot be factored out from polynomials, so, as for Dirac’s quantization procedure [22, 23, 24, 25], some care is required.
  • The physically observable momentum can be defined, as in (15), as 12gμνpν, with commutation relations of the form (27). This operator can be transformed, as for the Newton-Wigner operator [29], to the form −i∂∂xμ by the Foldy-Wouthuysen transformation [30] gx14pμgx−14.
  • We follow here essentially the method discussed in Reed and Simon [31] in their discussion of the Lebesgue integral.
  • Similar to the method followed in the simpler case of constant curvature by Georgiev [33].
  • Note that Abraham et al. [34] apply the formal Fourier transform on a manifold in three dimensions without proof.
  • Yahalom [36] has proposed an alternative view involving the retardation effects associated with gravitational waves, presently being tested and developed. We do not discuss this approach further here.

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Lawrence P. Horwitz (July 18th 2019). Stueckelberg-Horwitz-Piron Canonical Quantum Theory in General Relativity and Bekenstein-Sanders Gauge Fields for TeVeS [Online First], IntechOpen, DOI: 10.5772/intechopen.88154. Available from:

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