Research Article | Open Access
From Generalized Dirac Equations to a Candidate for Dark Energy
We consider extensions of the Dirac equation with mass terms and . The corresponding Hamiltonians are Hermitian and pseudo-Hermitian ( Hermitian), respectively. The fundamental spinor solutions for all generalized Dirac equations are found in the helicity basis and brought into concise analytic form. We postulate that the time-ordered product of field operators should yield the Feynman propagator ( prescription), and we also postulate that the tardyonic as well as tachyonic Dirac equations should have a smooth massless limit. These postulates lead to sum rules that connect the form of the fundamental field anticommutators with the tensor sums of the fundamental plane-wave eigenspinors and the projectors over positive-energy and negative-energy states. In the massless case, the sum rules are fulfilled by two egregiously simple, distinguished functional forms. The first sum rule remains valid in the case of a tardyonic theory and leads to the canonical massive Dirac field. The second sum rule is valid for a tachyonic mass term and leads to a natural suppression of the right-handed helicity states for tachyonic particles and left-handed helicity states for tachyonic spin-1/2 antiparticles. When applied to neutrinos, the theory contains a free tachyonic mass parameter. Tachyons are known to be repulsed by gravity. We discuss a possible role of a tachyonic neutrino as a contribution to the accelerated expansion of the Universe “dark energy.”
1.1. Generalized Dirac Equations: Mass Terms and Dispersion Relations
Dirac is often quoted saying in some of his talks that the equation that carries his name [1, 2] is “more intelligent than its inventor.” Of course, it needs to be added that it was Dirac himself who found most of the additional insight. Here, we are concerned with extensions of the Dirac equation which contain both tardyonic and tachyonic mass terms. Tardyonic (subluminal) mass terms lead to dispersion relations of the form , whereas tachyonic mass terms lead to superluminal dispersion relations of the form , where is the energy and is the momentum. The generalized, matrix-valued mass term enters the Dirac equation in the form . The are matrices that fulfill the relations , where we choose the space-time metric as . The denotes the partial derivative with respect to the space-time coordinate . It is quite surprising that a systematic presentation of the solutions of the generalized Dirac equations , in the helicity basis , has not been recorded in the literature to the best of our knowledge. While the following discussion is somewhat technical, we believe that it will be beneficial to give their explicit form, in order to fix ideas for the following discussion.
We set and use the Dirac matrices in the standard representation and define . For the ordinary Dirac theory, one has (one should say more precisely ) with a real mass , The dispersion relation is . The corresponding Dirac Hamiltonian reads Extensions of the Dirac equation with pseudoscalar mass terms that contain the fifth current have been introduced in the literature. In , it is shown that for a mass term of the form , the fermion propagator may obtain nontrivial gradient corrections already at the first order in derivative expansion, for a position-dependent mass. In that case, the fermion self-energy may contribute to a conceivable explanation for -violation during electroweak baryogenesis, as pointed out in . We thus study the following generalized form of the tardyonic (subluminal) Dirac equation: The dispersion relation is . The Hermitian tardyonic Hamiltonian operator reads as We may indicate a further motivation for our study; namely, the unitarity of the matrix implies the existence of useful relations  for the even powers obtained upon expanding a one-loop amplitude, formulated with a mass term , in powers of . This implies that a better understanding of the tardyonic equation with two mass terms could be of much more general interest.
It has not escaped our attention that the chiral transformation connects the two Hamiltonians and for and , but it is computationally easier and more instructive to consider the real and imaginary parts of the mass term separately.
Within a systematic approach to generalized Dirac equations with pseudoscalar mass terms, we also consider tachyonic (superluminal) mass terms of the form which induce a superluminal dispersion relation . The corresponding generalized Dirac equation has been named the “tachyonic Dirac equation” and reads as follows [6–10]: The corresponding Hamiltonian reads The relation has been given in [9, 10]. However, it is much more instructive to observe that is Hermitian; that is, . The concept of Hermiticity is known in lattice theory [11, 12] and is otherwise called pseudo-Hermiticity [13–23].
An obvious generalization of the tachyonic case contains an imaginary mass and a mass term, The dispersion relation is . The corresponding Hamiltonian reads and is Hermitian, . For , (9) has been discussed in [24, 25].
It is our goal here to present the fundamental eigenspinors corresponding to the plane-wave solutions of (2), (4), (7), and (9) in a unified and systematic manner. Furthermore, we discuss the second-quantized versions of the fermionic theories described by the generalized Dirac equations. Anticipating part of the results, we may point out that the massless Dirac equation “interpolates” between the tardyonic equations (2) and (4) and the tachyonic equations (7) and (9). For zero mass, helicity and chirality are equal. Helicity and chirality “depart” from each other in very specific directions, when the tardyonic and tachyonic mass terms are “switched on,” as we shall discuss in the following.
1.2. Tachyonic Dirac Equation and Neutrinos: Possible Connections
The tachyonic generalized Dirac equations (7) and (9) describe the motion of superluminal particles, which may either be important for astrophysical studies (neutrinos) or for artificially generated environments such as honeycomb photonic lattices in which pertinent dispersion relations become practically important . The existence of superluminal particles would not falsify Einstein’s theory of special relativity , which according to common wisdom is based on the following postulates. (i) The principle of relativity states that the laws of physics are the same for all observers in uniform motion relative to one another. (ii) The speed of light in a vacuum is the same for all observers, regardless of their relative motion or of the motion of the source of the light. Predictions of relativity theory regarding the relativity of simultaneity, time dilation, and length contraction would not change if superluminal particles did exist. Furthermore, as shown by Sudarshan et al. [28–31] and Feinberg [32, 33], the existence of tachyons, which are superluminal particles fulfilling a Lorentz-invariant dispersion relation , is fully compatible with special relativity and Lorentz invariance. According to special relativity, it is forbidden to accelerate a particle “through” the light barrier (because for ), but a genuinely superluminal particle remains superluminal upon Lorentz transformation. Significant problems are encountered when one attempts to quantize the tachyonic theories, but again, as shown in , these problems may not be as serious as previously thought. In particular, the so-called reinterpretation of solutions propagating into the past according to the Feynman prescription  is a cornerstone of modern field theory. Furthermore, it has been shown in  that tachyonic particles can be localized and equal-time anticommutators of the spin-1/2 tachyonic field involve an unfiltered Dirac- (see equation (37) of ).
Despite these arguments, we can say that, from the point of view of fundamental symmetries, accepting a superluminal neutrino would be equivalent to “an ugly duckling.” Adding to the difficulties, we notice that recent experimental claims regarding the conceivable observation of highly superluminal neutrinos have turned out to be false. One may point out that a relative deviation with at , as claimed by some recent experimental collaborations, would correspond to a negative neutrino mass square in the order of , if one assumes a Lorentz-invariant dispersion relation . Still, there is at present no conclusive answer regarding the conceivable superluminality of at least one neutrino flavor [35–39], and it is intriguing that all available direct measurements of the neutrino mass square have resulted in negative expectation values, still compatible with zero within experimental uncertainty, whereas published experimental best estimates for the neutrino speed [40–43] have been superluminal, again still compatible with the speed of light within experimental error. The recent ICARUS result  is consistent with this trend [44–50]; the best estimate for the neutrino velocity is superluminal, but the deviation from is statistically insignificant. The OPERA collaboration  has indicated a preliminary, revised result of . Neither subluminal nor superluminal propagation velocities are excluded based on the available experimental data. The “ugly duckling of a superluminal neutrino” is not beautiful; if we are to consider accepting it, then we should be able to hope that the emergence of at least one “added insight” should be the result of this operation.
Before we discuss the possible emergence of these benefits, let us include some historic remarks. According to reliable sources (Professor M. Fink from the University of Austin (Texas) was engaged in discussions with Professor J. A. Wheeler, who found the notion of a nonvanishing neutrino mass so unappealing that he discouraged experimentalists from undertaking any effort to measure the neutrino mass, based on arguments described in Section 1 (M. Fink, private communication, 2012.)), Professor J. A. Wheeler, in his later years at the University of Austin (Texas), used to argue that the neutrino has to be massless, necessarily, and that in his opinion, it could only be a massless Weyl particle with definite helicity (and chirality). Recently presented arguments  regarding the possibility of overtaking a subluminal, left-handed neutrino, looking back and seeing a right-handed neutrino, were supposedly already used by Wheeler in order to dispel the conceivable existence of a neutrino mass term. If we assume that right-handed neutrinos are much more massive than their left-handed counterparts, then this leads to a paradox unless one assumes that all light, right-handed neutrinos are sterile. (The problem with a right-handed sterile massive neutrino is that for massive neutrinos, chirality and helicity are different; hence a coupling of the form no longer vanishes for massive Dirac neutrinos if one uses the canonical eigenstates of the massive Dirac equation. One therefore has to invoke additional exotic mechanisms in order to ensure the “sterility” of the right-handed Dirac neutrinos.) Wheeler also disliked (Professor M. Fink from the University of Austin (Texas) was engaged in discussions with Professor J. A. Wheeler, who found the notion of a nonvanishing neutrino mass so unappealing that he discouraged experimentalists from undertaking any effort to measure the neutrino mass, based on arguments described in Section 1 (M. Fink, private communication, 2012.)) the notion of a Majorana neutrino, arguing that the charge conjugation invariance condition imposed on the Majorana particle precludes the existence of plane-wave solutions to the Majorana equation and maximally violates lepton number. Again, these arguments (We are grateful to Professor M. Fink for the clarification and confirmation.) are in full agreement with those recently given in .
The original standard model thus called for manifestly massless neutrinos. The commonly accepted observation of neutrino oscillations precludes the possibility that all three generations of neutrino mass eigenstates are massless. Lepton number conservation is based on the global gauge symmetry , applied simultaneously to all lepton fields. A Majorana neutrino would destroy lepton number as a global symmetry but solve the “autobahn paradox,” because a Majorana neutrino would be equal to its own antiparticle and thus, looking back, the right-handed neutrino state would consist of the same particle = antiparticle.
On the other hand, if we assume that the neutrino is described by the tachyonic Dirac equation, then the following statements are valid.(i)Statement 1: we can properly assign lepton number and use plane-wave eigenstates for incoming and outgoing particles, while allowing for nonvanishing mass terms and thus mass square differences among the neutrino mass (not flavor) eigenstates. (ii)Statement 2: there is a natural resolution for the “autobahn paradox” because a left-handed spacelike neutrino always remains spacelike upon Lorentz transformation and cannot be overtaken. (iii)Statement 3: the right-handed particle and left-handed antiparticle states are suppressed due to negative Fock-space norm.(iv)Statement 4: at least qualitatively, tachyonic neutrinos could yield an explanation for a repulsive force on intergalactic distance scales as they are repulsed, like all tachyons, by gravitational interactions (dark energy).
Pauli  postulated the existence of neutrinos, on the basis of the conservation of angular momentum and energy, and also introduced pseudo-Hermitian operators . Here, we describe conceivable connections of neutrino physics and pseudo-Hermitian operators. Final clarification can only come from experiment. When in 1956, Reines and Cowan  discovered the electron neutrino, two and a half years before Pauli’s death, Pauli replied  by telegram “Thanks for message. Everything comes to him who knows how to wait. Pauli.” In defense of the tachyonic hypothesis, we would like to stress that a tachyonic Dirac neutrino would allow us to retain lepton number conservation as a symmetry of nature. We would thus like to write up these thoughts in the current paper, with attention to detail. We should point out that our approach fully conserves Lorentz invariance, in contrast to the extensions of the Standard Model based on Lorentz-violating terms which can otherwise lead to superluminal propagation (see Tables 11 and 13 of ).
Units with are used throughout the paper. The organization is as follows: in Section 2, we discuss massless and tardyonic theories. Tachyonic extensions of the Dirac equation are discussed in Section 3. Connections of the fundamental tensor sums over the eigenspinors with the derivation of the time-ordered propagator are analyzed in Section 4. A candidate for dark energy is presented in Section 5. Conclusions are reserved for Section 6.
2. Generalized Dirac Equations: Massless and Tardyonic Theories
2.1. Massless Dirac Theory
The massless Dirac equation and the massless Dirac Hamiltonian read as We note that is both Hermitian as well as Hermitian; that is, . The dispersion relation is . With , we seek positive-energy and negative-energy solutions of the form where denotes a quantum number which is equal to the helicity for positive-energy states and equal to the negative of the helicity for negative-energy states. With , we have . In the massless limit, the solutions to the Dirac equation are given as (see Chapter 2 of )The well-known helicity spinors are recalled as These fulfill the fundamental relations , as well as and , where and the sum over is over the values . The sums over the fundamental bispinors and fulfill the following sum rules: as well as Sum rule I can be obtained by a quick explicit calculation, and sum rule II holds because in the massless limit, helicity equals chirality (positive sign for positive energy, negative sign for negative energy). We denote the Dirac adjoint by . One can easily check by an explicit calculation that and . We can thus introduce a factor under the summation over spins in (15) and replace one of the factors by a multiplication of the Dirac adjoint spinor from the right by the fifth current. The Lorentz-invariant normalization of the massless solutions vanishes; that is, .
Eigenstates of the massless Hamiltonian have to be eigenstates of the chirality operator because the chirality commutes with the Hamiltonian, in the sense that . Furthermore, where is the vector of spin matrices, and the helicity operator is identified as . Let be the eigenvalue of chirality and let be the eigenvalue of the helicity operator. Then, the eigenvalue of the Hamiltonian is . Since and , we easily recover the known fact that helicity equals chirality for positive energy, whereas the relation is reversed for negative-energy states (see also Chapter 2.4 of ). We are aware of the fact that the considerations reported in the current section partly refer to the literature, but we give them in some detail because they are essential for the following considerations.
2.2. Massive Dirac Theory
We start from the ordinary Dirac equation given in (2), which reads . In the helicity basis, the fundamental spinor solutions read as The algebraic relations that have to be fulfilled by the bispinor amplitudes and read as follows: The dispersion relation is . In the helicity basis, the solutions of (19) with a tardyonic mass term are easily written down, using the identity . With an appropriate normalization factor and after some algebraic simplification, the positive-energy solutions read as follows:The negative-energy eigenstates of the tardyonic equations in the helicity basis are given asThese solutions are consistent with those given in  and in Chapter 23 of , and the normalizations are () One can change the normalization according to The Lorentz-invariant normalization is equal to one for the fundamental positive-energy bispinors and equal to minus one for the fundamental negative-energy bispinors, A little algebra is sufficient to reproduce the following known sums over bispinors: In accordance with general wisdom of the tardyonic case, these do not involve a helicity-dependent prefactor. The sum rule (25) is of type I (see (15)).
2.3. Two Tardyonic Mass Terms
Inspired by the discussion in Section 1, we consider an equation with two tardyonic mass terms, which has already been indicated in (2) and reads . For the corresponding bispinors in the fundamental plane-wave solutions, this implies thatThe dispersion relation is . The fundamental positive-energy bispinors read as follows:The negative-energy eigenstates of the equation with two tardyonic mass terms are given as for negative helicity (positive chirality in the massless limit) and for positive helicity (negative chirality in the massless limit). In the massless limit (first , then , and then ), we again reproduce the massless solutions , , and . Of course, in the limit , one also has to expand the normalization denominators in powers of . For the solutions (27a), (27b), (28a), and (28b) reduce to the solutions of the ordinary Dirac equation in (20a), (20b), (21a), and (21b) and can be expanded in to yield corrections to the ordinary Dirac equation for small ; that is, . The states are normalized with respect to the condition In the normalization the positive-energy solutions acquire a “positive Lorentz-invariant norm,” whereas the negative-energy solutions have “negative Lorentz-invariant norm,” After some algebra, one can derive the following sums over bispinors: These are easily identified as the positive- and negative-energy projectors. The sum rules do not involve helicity-dependent prefactor and are of type I (see (15)).
The solutions (27a), (27b), (28a), and (28b) approach the massless solutions if one replaces first and then let . They are thus useful for systems where the mass is greater than . For , one would like to calculate solutions that approach the massless case for the sequence , then . These read as follows: In comparison to the solutions (27a), (27b), (28a), and (28b), they acquire a nontrivial phase factor.
3. Generalized Dirac Equations: Tachyonic Mass Terms
3.1. Tachyonic Dirac Equation
The tachyonic Dirac equation is given in (7) and reads . The fundamental bispinors entering the equations fulfill the equations Using and some algebra, the prefactors in the fundamental bispinors (for positive energy) take a very simple form,For negative energy, the solutions read as follows: The normalization condition is () One can change to Lorentz-invariant normalization by a multiplication with , The “calligraphic” spinors fulfill the following helicity-dependent normalizations: where we observe that is a good quantum number because the helicity operator commutes with the Hamiltonian (8). The sum rule fulfilled by the fundamental plane-wave spinors is of type II (see (16)). For the positive-energy spinors, we have where for the negative-energy spinors, the sum rule reads The expressions on the right-hand sides are the positive- and negative-energy projectors.
3.2. Two Tachyonic Mass Terms
We study the equation , as given in (9). The fundamental spinors, which we denote and , fulfill the following equations:
The positive-energy solutions are obtained using the identity . With , they read as follows:
The negative-energy solutions for the tachyonic equation with two mass terms are given asThe normalization condition is . We use a definition of the “calligraphic” spinors analogous to (38), In analogy to (40a) and (40b), a sum rule of type II (see (16)) is fulfilled by the fundamental plane-wave spinors, We thus obtain that the desired projectors onto positive- and negative-energy solutions for the Dirac equation with two tachyonic mass terms (9). The generalized equation is fully compatible with the Clifford-algebra-based approach recently described in .
4. Theorems for Generalized Dirac Fields
4.1. Spinor Sums and Time-Ordered Propagator
Our central postulate regarding the quantized fermionic theory is that the time-ordered vacuum expectation value of the field operators should yield the time-ordered (Feynman) propagator, which, in the momentum representation, is equal to the inverse of the Hamiltonian (upon multiplication with ). This postulate implies that under rather general assumptions regarding the mathematical form of the elementary field anticommutators, sum rules have to be fulfilled by the tensor sums over the fundamental spinor solutions. It is perhaps not surprising that these sum rules are precisely of the form investigated in Sections 2 and 3 of this paper.
For definiteness, we consider the solution of the tachyonic Dirac equation (Section 3.1). The generalization to other generalized Dirac equations is straightforward. We start from the field operator  where annihilates particles and creates antiparticles. The following anticommutators vanish:We assume the following general form for the nonvanishing anticommutators,with arbitrary and functions of the quantum numbers and . One might argue that since and must be dimensionless, they can depend only on the dimensionless arguments and , but that is a detail of the discussion which we do not pursue any further. Our only assumption concerns the fact that the field anticommutators should be diagonal in the helicity and wave vector quantum numbers, leading to the corresponding Kronecker and Dirac-’s.
We assume that the spin-matrix either constitutes a Lorentz scalar or a pseudoscalar quantity, which is a scalar under the proper orthochronous Lorentz group. The time-ordered product of field operators reads as This equation contains the same coefficient functions and that enter into (48a) and (48b). In order to proceed with the derivation of the propagator, we must postulate that the following sum rules hold: The sum rule (50) is crucial for the further steps in the derivation of the time-ordered propagator. Introducing a suitable complex integral representation for the step function, one obtains from (49), using (50), after a few steps which we do not discuss in further detail (see also equations (3.169) and (3.170) of ) The convention is that in any integrals , the component is set equal to in the integrand when it occurs in scalar products of the form , and so forth, but if the integral is over the full , then the integration interval is the full . With the convention which is adopted in many quantum field theoretical textbooks, including [57, 60], we finally obtain the result The tachyonic propagator is identified, under the integral sign, as The sum rules (40a) and (40b) imply that the derivation is valid for the choice in which case the relations given in (50) are fulfilled. Note that this observation does not imply that the choice (54) necessarily is the only one for which we are able to fulfill the postulates given in (50), but it is the only structurally simple choice that we have found.
For the egregiously simple choice (54), let us study the transition to the massless limit (16) in some further detail. Indeed, in the limit , the denominator of the spin sums in (40a) and (40b) vanishes, and a finite limit is obtained after multiplication with ,In order to compare the normalizations of the fundamental spinors in the massless limit, we calculate the following quantities: where the fundamental spinors and of the massless equation have been given in Section 2.1. Observing that in the massless limit, the identifications and , as implied by (56) and (57), show that the identities (55a) and (55b) precisely reduce to the sum rule (16) in the massless limit.
4.2. Generalized Field Anticommutators for Tardyonic and Tachyonic Fields
For definiteness, we have considered the case of the tachyonic Dirac field in the above derivation. The decisive observation is that the tachyonic choice, is consistent with both massive tachyonic fields discussed in Sections 3.1 and 3.2, whereas the tardyonic choice, yields the time-ordered propagator for both massive tardyonic fields discussed in Sections 2.2 and 2.3. The nonvanishing anticommutators for tardyons take the simple form (cf. (48a) and (48b)), Again, compared with (48a) and (48b), the nonvanishing anticommutators for tachyons take the simple form, With these universal choices, the theory of the tardyonic and tachyonic spin-1/2 fields can be unified. The time-ordered propagator is given as We use the sum rules for the tensor sums over fundamental spinors given in (25) (for the tardyonic Dirac field), (32) (for the tardyonic Dirac field with two mass terms), (40a) and (40b) (for the tachyonic Dirac field), and (45) (for the Dirac field of imaginary mass). Going through the exact same derivation as outlined above in between (49) and (53), we obtain the following results for the time-ordered propagators of tardyonic and tachyonic fields. For the tardyonic Dirac field (Section 2.2), one has For the tardyonic field with two mass terms (Section 2.3), the Feynman propagator is easily found as where and are infinitesimal imaginary parts. Both tardyonic mass terms acquire an infinitesimal negative imaginary part, and the prefactor from (46) in the field operator needs to be replaced by , where the tardyonic energy is . For the tachyonic Dirac field (Section 3.1 in ), one has the result given in (53). Finally, for the Dirac field with two tachyonic mass terms, we have In the latter case, the prefactor from (46) in the field operator needs to be replaced by , where the tachyonic energy is . For both tachyonic fields discussed here, the mass acquires an infinitesimal positive imaginary part, as manifest in the results given in (53) and (65).
4.3. Tachyonic Gordon Identities
It is useful to illustrate the derivation outlined above by exploring its connection to tachyonic Gordon identities. For definiteness, we again concentrate on the tachyonic Dirac equation discussed in Section 3.1. The matrix element of the vector current finds the following Gordon decomposition for positive-energy spinors: For negative-energy solutions, the identity reads as For , one hasThe matrix element of the axial current reads whereas for negative-energy solutions For , this simplifies toThe results (69a), (69b), (70a), and (70b) for the tachyonic axial vector current have a similar structure as the Gordon decomposition for the tardyonic vector current obtained with the ordinary Dirac equation (see equation of ). The role of the Dirac adjoint for the tardyonic case is taken over by the “chiral adjoint” for the tachyonic particle. Here, the designation “chiral adjoint” is inspired by the fact that transforms as a pseudoscalar under Lorentz transformations.
The structure of (68a), (68b), (70a), and (70b) is somewhat peculiar with regard to parity. In (68a), an apparent vector current on the left-hand side appears to transform into an axial current on the right-hand side, whereas in (70a), an apparent axial vector on the left-hand side of the equation becomes what appears to be a vector on the right-hand side. The reason lies in the more complicated behavior of the tachyonic Dirac equation under parity as investigated in . Namely, the tachyonic Dirac equation (7) contains a term which transforms as a scalar under parity, as well as a term which transforms as a pseudoscalar, The mass term in the tachyonic Dirac equation is pseudoscalar and changes sign under parity. Indeed, in , the tachyonic Dirac equation has been shown to be separately invariant, and invariant, but not invariant, due to the change in the mass term.
In order to put this observation into perspective, we recall that the entries of the electromagnetic field strength tensor are composed of axial vector components (magnetic field), as well as vector components (electric field). The transformation properties of the electromagnetic field strength tensor under the proper orthochronous Lorentz group are nevertheless well defined.
The transformation (71b) can be interpreted as a transformation under parity. Thus, if we interpret the mass as a pseudoscalar quantity, then the right-hand sides of (68a) and (70a) transform as a vector and an axial vector, respectively. It is the parity noninvariance of the mass term in the tachyonic Dirac equation which leads to the somewhat peculiar structure of (68a) and (70a).
4.4. Helicity-Dependence and Gupta-Bleuler Condition
The anticommutator relations for tardyons given in (60) imply that both left-handed and right-handed helicity states, for both particles and antiparticles, have positive norm. However, the anticommutator relations for tachyons given in (61) imply that right-handed particle as well as left-handed antiparticle states acquire negative norm. This is shown in equations (31) and (32) of . Indeed, for one-particle states , where is the normalization volume in coordinate space. The Fock-space norm is negative for .
For clarification, the corresponding Gupta-Bleuler condition should be indicated explicitly. In full analogy to the instructive discussion of the Gupta-Bleuler mechanism for the photon field, as given in full clarity in Chapter (9b) of Schweber’s textbook , we select the positive- and negative-frequency component of the ()-component of the neutrino field operator,and postulate that it annihilates any physical Fock state of the tachyonic field, These relations automatically imply that the Gupta-Bleuler condition also is realized in terms of the expectation value but the condition (74) is stronger. We recall that the Gupta-Bleuler condition on the photon field reads , where is a Fock state of the photon field. As stressed in Schweber’s book  on page 246, the condition is not sufficient for the suppression of the longitudinal and scalar photons, but one must postulate that , where is the positive-frequency component of the photon field operator. Because the tachyonic fermion, unlike the photon, is not equal to its own antiparticle, we need two conditions, given in (74).
A crucial question now concerns the possibility of reversing the helicity-dependence, that is, the question of whether or not a different choice for the helicity-dependent factors in (48a) and (48b) exists that would imply negative norm for left-handed particles and right-handed antiparticles. A related question is whether other tachyonic Dirac Hamiltonians exist for which the choice instead of would fulfill our general postulate, namely, the sum rule (50). For reasons outlined in the following, we can ascertain that this is not the case; tachyonic spin-1/2 particles should always be left handed.
The arguments supporting this conclusion are as follows. First, if we assume that the tachyonic field fulfills a sum rule of type II which for massless fields is given in (16), then it is impossible to replace by in the sum rule because of the necessity to preserve a smooth massless limit. The considerations in the text following (16) imply that if one were to replace by in the massless case, then one would violate the sum rule (16). The second argument is obtained by explicit calculation. We have checked that if one replaces in the tachyonic and imaginary-mass Dirac equations (7) and (9), then the sum rules fulfilled by the corresponding fundamental spinors still contain the characteristic factor . For the imaginary-mass Dirac equation, this result is obtained in . Intuitively, we can understand this result as follows: the mass in the denominators of the right-hand sides of (40a), (40b), and (45) is obtained as the modulus and does not change if we replace in the superluminal Dirac equation. The mass in the numerator of the right-hand sides of (40a), (40b), and (45) changes sign, but this is consistent with the obvious change in the functional form of the positive-energy and negative-energy projectors as we change the sign of the mass term. Again, this consideration supports the conclusion that we cannot invert the helicity-dependence by choosing a different Hamiltonian; the factor persists.
The third argument comes from the tachyonic Gordon identities discussed in Section 4.3. We use the Gordon identity (70a) and (70b) and the normalization (39) to calculate the bispinor trace (with a multiplied from the right) of the left-hand side of (40a), The bispinor trace of the right-hand side of (40a) is which shows the consistency of the bispinor sum (40a) with the Gordon decomposition (70a) and (70b). If we were to replace by , the two sides of the relation (76a) would differ by a minus sign. The bispinor trace of the left-hand side of (40b) is We have used the tachyonic Gordon decomposition for negative-energy states as given in (70a) and (70b), which differs from the positive-energy Gordon decomposition by a minus sign, but an additional minus sign is obtained from the Lorentz-invariant normalization of the negative-energy fundamental bispinors. From the right-hand side of (40b), we have