Article General Quantum Field Theory of Flavor Mixing and Oscillations
Chueng-Ryong Ji 1,* and Yuriy Mishchenko 2
1 Department of Physics, North Carolina State University, Raleigh, NC 27695-8202, USA 2 Amazon.com Inc., 101 Main St., Cambridge, MA 02142, USA; [email protected] * Correspondence: [email protected]
Abstract: We review the canonical transformation in quantum physics known as the Bogoliubov transformation and present its application to the general theory of quantum field mixing and oscillations with an arbitrary number of mixed particles with either boson or fermion statistics. The mixing relations for quantum states are derived directly from the definition of mixing for quantum fields and the unitary inequivalence of the Fock space of energy and flavor eigenstates is shown by a straightforward algebraic method. The time dynamics of the interacting fields is then explicitly solved and the flavor oscillation formulas are derived in a unified general formulation with emphasis on antiparticle content and effect introduced by nontrivial flavor vacuum.
Keywords: quantum field theory; flavor mixing and oscillations; nontrivial vacuum
1. Introduction The importance of the proper choice of the degrees of freedom can hardly be un- Citation: Ji, C.-R.; Mishchenko, Y. derestimated in physics. Taking the form of canonical transformations and generalized General Quantum Field Theory of coordinates formalism in classical mechanics, the transformations of degrees of freedom Flavor Mixing and Oscillations. had been taken to a qualitatively different level in quantum physics and are responsible Universe 2021, 7, 51. https:// for such dramatic phenomena as phase-transitions, superconductivity, superfluidity and doi.org/10.3390/universe7030051 quark-hadron duality. Representations of physical theory in terms of different generalized coordinates in classical mechanics were emphasized and elaborated in the theory of canon- Academic Editor: Giuseppe Gaetano ical transformation (CT) using Hamiltonian dynamics and the Jacobi Equation. The theory Luciano of CT plays a central role in classical mechanics both by providing a powerful tool for solving dynamics and by setting up a conceptual framework to establish bridges between Received: 15 January 2021 Accepted: 19 February 2021 various physical models. However, the general theory of CT fails in quantum mechanics Published: 28 February 2021 due to the operator nature of the generalized coordinates that leads to the ordering ambi- guities in generating functions. Although some conjectures existed that identified CT in
Publisher’s Note: MDPI stays neutral quantum mechanics with unitary transformations and even some classes of non-unitary with regard to jurisdictional claims in transformations, the true relevance of CT in quantum physics is yet to be established. Out published maps and institutional affil- of a large body of CT known in classical mechanics, the only transformation that survived iations. the quantum transition is what is known today as the Bogoliubov Transformation (BT). In this work, we consider the role of BT in the field theory of mixing and the definition of the flavor quantum states and make an attempt toward a complete formulation of such a field theory of mixing. This involves an introduction of general formalism as well as
Copyright: © 2021 by the authors. phenomenological implications. Quantum mixing of particles is among the most interest- Licensee MDPI, Basel, Switzerland. ing and important topics in particle physics [1–6]. The standard model involves quantum This article is an open access article mixing in the form of the Kobayashi–Maskawa (CKM) mixing matrix [7], a generalization distributed under the terms and of the original Cabibbo mixing between d and s quarks [8]. Also, recently, convincing conditions of the Creative Commons evidence of neutrino mixing have been provided by Super-Kamiokande and SNO experi- Attribution (CC BY) license (https:// ments [5,6,9–14], thus suggesting neutrino oscillations as the most likely resolution for the creativecommons.org/licenses/by/ solar neutrino puzzle [15–18] and the neutrino masses [19–22]. 4.0/).
Universe 2021, 7, 51. https://doi.org/10.3390/universe7030051 https://www.mdpi.com/journal/universe Universe 2021, 7, 51 2 of 27
Since the middle of the century, when the quantum mixing was first observed in meson systems, this phenomenon has played a significant role in the phenomenology of particle physics. Back in the 1960s, the mixing of K0 and K¯ 0 provided evidence of CP-violation in weak interactions [23] and till today meson mixing is used immensely to experimentally determine the precise profile of the CKM unitarity triangle [7,8,24]. Upgraded high-precision B-meson experiments would be vital to search for deviations from the unitarity in a CKM matrix, which can put important constraints on the physics beyond the standard model. At the same time, in the fermion sector, the discovery of neutrino mixing and neutrino masses challenged our fundamental understanding of the P-violation and, in part, of the standard model itself. Regarding the vanishing magnitudes of the expected new physics effects (such as the unitarity violation in the CKM matrix and/or neutrino masses), it is imperative that the theoretical aspects of the quantum mixing are precisely understood. In this respect, it was noticed recently that the conventional treatment of flavor mixing, where the flavor states are defined in the Fock space of the energy eigenstates, suffers from the problem with the probability conservation [25,26]. This suggested that the mixed states should be treated independently from the energy eigenstates. It was found, indeed, that the flavor mixing in quantum field theory introduces very nontrivial relationships between the flavor and the energy quantum states, which lead to unitary inequivalence between the Fock space of the interacting fields and that of the free fields [27–30]. This is quite different from the conventional perturbation theory where one expects the vacuum of the interacting theory to be essentially the same as one of the free theory (up to a phase factor eiS0 [31,32]). Recently, the importance of mixing transformations has prompted their fundamental examination from a field-theoretical perspective. The investigation of two-field unitary mixing in the fermion sector by Blasone and Vitiello [25,30,33,34] demonstrated a rich structure of the interacting-field vacuum as a SU(2) coherent state and altered the oscilla- tion formula to include the antiparticle degrees of freedom. Subsequent analysis of the boson case revealed a similar but much richer structure of the vacuum of the interacting fields [35,36]. Especially, the pole structure in the inner product between the vacuum of the free theory and the vacuum of the interacting theory was found and related to the convergence radius of the perturbation series [36]. Attempts to look at the mixing of more than two flavors have also been carried out [37,38]. Also, a mathematically rigorous study of 2-flavor quantum field mixing has been offered by Hannabuss and Latimer [39,40]. In this article, we present a brief review of CT and apply the quantum linear CT (or BT) to the advances in quantum field theory of mixing. The mixing relations for quantum states here are derived directly from the definition of the mixing for quantum fields and the unitary inequivalence of the Fock spaces of energy and flavor eigenstates is shown by a direct algebraic method. The time dynamics of the interacting fields is explicitly solved and the flavor oscillation formulas are derived in a general form with emphasis on antiparticle content and the effect from the nontrivial flavor vacuum. The paper is organized as follows. In Section2, we discuss the theory of CT with a special emphasis on quantum linear CT (or BT) as they appear in classical and quantum physics discussing their novel applications even in superfluidity and low-energy QCD. In Section3, we analyze quantum mixing from a general perspective of quantum field theory and introduce most general field-theoretical linear mixing relations. We then analyze implications of these relations for the flavor vacuum state, flavor Fock space and time dynamics of flavor states in quantum field theory. In Section4, we present the mixing matrices and discuss the mass parameterization issue. We obtain the mass parameters in rather general form for boson/fermion mixing with an arbitrary number of flavors and discuss the issue of the mass parametrization. Conclusions follow in Section5. Supplemental materials are placed in the Appendices. AppendixA is devoted to explicitly solving the structure of the nontrivial flavor vacuum state in the general quantum field theory of flavor mixing. In AppendixB, we list explicitly the mixing parameters for Universe 2021, 7, 51 3 of 27
the most important cases of particle mixing. In AppendixC, we present the results of application of our general formalism to 2-flavor mixing for spin 0, 1/2 and 1 particles.
2. Canonical Transformations in Quantum Physics The theory of CT is a powerful tool of classical mechanics. Yet, the strength of CT is yet to be fully realized in quantum mechanics [41–44]. The ground for CT in classical mechanics is laid down by the Hamilton formalism in which the Lagrangian function and the dynamics in terms of generalized coordinates
∂ ∂ L(q, q˙, t) − L(q, q˙, t) = 0 (1) ∂q˙i ∂qi
is translated into dynamics in terms of twice as many coordinates and momenta and the Hamiltonian function ∂ q˙i = H(q, p, t), ∂pi (2) p˙ = − ∂ H(q, p, t). i ∂qi Hamiltonian formalism is advantageous over Lagrangian formalism in that it reduces the system’s dynamics to the solution of a set of first order differential equations. On the other hand, one has to deal with twice as many variables. The main advantage of the Hamiltonian formalism is, however, in the conceptual framework. Note that in Lagrangian formalism the only “fundamental” degrees of freedom are the generalized ∂L coordinates and the momenta are merely derivatives p = ∂q˙ . In Hamiltonian formalism both coordinates and momenta are treated on equal footing and are independent except for the dynamical link established by Hamilton’s Equation (2). Such independence provides the most important piece of the foundation for the theory of canonical transformations. Thus, in classical mechanics CT is such a change of the phase space variables (q, p) → (Q, P) that preserves the Poisson bracket
[q, p] = 1 = [Q, P]. (3)
The Poisson bracket is defined as usual,
∂ f ∂g ∂ f ∂g [ f , g] = − . (4) ∂q ∂p ∂p ∂q
The main property of CT is the way the action is transformed Z Z Z dt(pq˙ − H(p, q, t)) = dF(P, Q, t) + dt(PQ˙ − H0(P, Q, t)), (5)
where the full differential is given by
dF = p · dq − P · dQ. (6)
As long as Equation (5) holds, the dynamics of the system in new coordinates is described with Hamilton formalism and, in this sense, is similar to that of the original description. F(P, Q, t) is often used to characterize classical CT and is typically called a generating function [45]. CTs are extremely helpful tools that allow to change the system’s Hamiltonian to a simpler form thus leading to great simplifications in the equations of motion. A textbook example in this respect is the oscillator dynamics, which can be transformed to a trivial problem with decoupled variables with [45]
P = √1 (ip + x) 2 (7) Q = √1 (−ip + x), 2 Universe 2021, 7, 51 4 of 27
conventionally denoted as a and a†. CT can also be used to generate families of exactly solvable Hamiltonians out of a single Hamiltonian where the dynamics are known, thus, providing a set of "toy" models in which the properties can be studied exactly. Finally, CT plays a central role in the Jacobi theory where a special transformation is sought that reduces the dynamics to a trivial one H → H0 = 0 and thus provides immediately the full solution to the classical equations of motion [45]. In quantum mechanics, however, the use of CTs is practically completely lost. Three major problems exist in the translation of the formalism of CTs to the quantum mechanics [46–48]: the ordering of operators must be specified, the inverse and fractional powers of operators that may appear in the transformation must be handled and the possibility of non-unitary CT must be addressed. While in classical mechanics three major roles of canonical transforma- tions (evolution, physical equivalence and solving theory) are blurred together, in quantum mechanics they are distinct. In quantum mechanics the evolution is produced by unitary transformations, while the physical equivalence is proved with isometric transformations (norm preserving isomorphisms between different Hilbert spaces) and the solution of a theory is achieved by general transformations, which may involve non-unitary transforma- tions [49,50]. A number of approaches had been pursued in a general attempt to resolve these issues [41–44,47,48,51–59]. The conjecture that CTs in quantum mechanics are one and the same with the unitary transformations is one of the oldest such attempts [51,52]. Non- hermitian linear transformations [56] and general form integral transformations [53–55] had also been considered in this respect. Such approaches typically experience the problem that the procedure necessary to build and apply a CT is not at all simpler than finding the solution of the original Schrodinger equation itself, thus undermining the very first idea of the use of CT for simplification of the original problem [51–55]. A rather different approach embraces the path integral as the base for further devel- opment of CT in the quantum theory [57–59]. The Feynman path integral (PI) written in the phase space of p and q provides one of the most startling paradoxes in the breakdown between classical and quantum CTs. Conventionally, one writes the PI in the Euclidean space as Z R U(q, q0; t) = Dq(t)Dp(t)e dt(ipq˙−H(p,q,t)). (8)
The similarity between Equations (5) and (8) may tempt one to use this relation as the basis for the program of quantization of CT [57–59]. Nonetheless it had been noted in the literature that the main obstacle along this line consists in the discretized nature of the expression (8) [44]. Specifically, one needs to remember that Equation (8) is only a formal representation in which the time derivatives, for example, should be properly defined. In this sense Equation (5) can not hold if p · dq − P · dQ 6= dF(P, Q) in a finite-difference form. Also, an explicit application of canonical transformation to formal expression (8) to derive, for example, the analog of the Jacobi equation for the propagator U(q, q0) immediately yields an inconsistent answer. The breakdown of Equation (6) in the quantum case can be seen as the primary source of the lack of correspondence between quantum and classical CTs. It may be of interest to examine this point more closely. Consider, for example, the point transformation q → Q(q) in the usual Lagrangian form of the path integral
R 2 U(q, q0; t) = R Dq(t)e− dt(q˙ /2+V(q)) → − R dt[( dQ )2Q˙ 2/2+V(Q)] (9) U(Q, Q0; t) = R DQ(t)J(Q)e dq .
This transformation is widely applied in the field theory and generally known to work well even though Equation (9) is only a formal representation of a properly discretized expression. One reason for this is that the contributions to the PI in Lagrangian form come only from the trajectories that are continuous, thus, justifying the use of transformation (9). Really, one considers a trajectory that has a discontinuity q(τ) → q(τ + 0) = q(τ) + ∆ at Universe 2021, 7, 51 5 of 27
some time τ. If sampled with time step dt, such trajectory would contribute to the integral a quantity − R dt(q˙2/2+V(q)) 2 e−∆ /2dte t6=τ . (10) As dt → 0, this contribution becomes exponentially suppressed relative to the contri- butions coming from the comparable continuous trajectories. Then, it can be said that the support of the integral in Equation (9) consists only from the continuous trajectories q(t) and this is why the formal operation with the integrand in Equation (9) works. In the case of the phase space form of PI (8), one may immediately observe that integra- tion over discontinuous trajectories (q(t), p(t)) is not suppressed. Really, the contribution of such a trajectory would come with merely a factor of ep∆q where ∆q is the discontinuity in q(t). Moreover, to derive the Lagrangian PI from Equation (9), one needs to integrate each p(x) from −∞ to +∞ regardless of the values of p(x) at the neighboring points. Thus, in Equation (8) the discretized nature of the integral is important and Equation (6), infinites- imally correct, cannot be generally used. It is useful to note, however, that Equation (6) will retain its general validity in finite-differenced form if the canonical transformation is linear. Indeed, it had been known for quite some time that linear canonical transformation can be applied successfully in quantum mechanics (most typical example is, again, quantum oscillator) [41–43]. Despite such severe limitations of the apparatus of CTs in quantum mechanics, the use of linear CT had proved to be of great advantage in the study of nonperturbative features of quantum systems. The quantum linear CT was first put forward in 1947 by Russian physicist N. Bogoliubov in order to build a microscopic theory of superfluidity and often bears his name [60]. The method was then extended to superconductivity in 1958 [61,62]. The theory of BT is usually formulated in terms of creation/annihilation operators of the quantum many-body problem. Two forms of BT (one for fermions and another for bosons) are known. For bosons BT reads
† Ak = u(k)ak − v(k)b−k † † (11) B−k = v(k)ak − u(k)b−k,
where u(k) and v(k) are transformation amplitudes such that u2(k) − v2(k) = 1. For fermions † Ak = u(k)ak + v(k)b−k † † (12) B−k = −v(k)ak + u(k)b−k and u2(k) + v2(k) = 1. The theory of BT can be viewed mathematically as a rotation of basis in the linear space of quantum fields built on (ak, bk, h.c.). BT is a unitary transformation and can be represented in the form † † i ∑ ρ(k)(ak b−k−akb−k) U = e k . (13) Physically, due to the manifestation of fundamental physical properties as observable effects inherent to quantum physics, this transformation describes an alternative set of degrees of freedom of the field-theoretic model that appear as quasi-particles with different properties than those of the original particles. Over the years, BT found a wide range of applications in various areas of quantum physics from condensed matter theory [63] to strongly interacting QCD [64–67]. It pro- vided a powerful nonperturbative tool that helps to understand many central features of macroscopic behavior of field-theoretic models. To illustrate the power of BT, let us begin by considering its first success in the theory of superfluidity in more detail. Superfluidity is the phenomenon of loss of viscous friction in a flowing fluid at superlow temperatures. Superfluidity was discovered by Petr Kapitza and Arno Allen Penzias and Robert Woodrow Wilson in 1938 [68]. P. Kapitza later received the Nobel prize for this discovery. The theoretical explanation of superfluidity was obtained in the works Universe 2021, 7, 51 6 of 27
of L. Landau and N. Bogoliubov [60,69,70]. According to Landau, viscous friction is due to the transfer of momentum between neighboring fluid elements toward the walls of the fluid container via the gradient in the fluid velocity. At low temperature, the viscous drag is transferred to the fluid from the stationary walls by means of elementary excitations. If such elementary excitation is created at the wall with momentum p, it will transfer to the fluid a momentum P0 = p and an energy E0 = e(p) so that the momentum and energy of the fluid become P = Mv + P0 1 2 (14) E = 2 Mv + E0 + P0v. The central point of this argument is the notion that the viscous drag excitations can be created at the walls only if there is free energy to do so, i.e., if the energy of the fluid will decrease when such excitation is created. That means that if
v < e(p)/p (15)
then none such elementary excitation can be created. For any fluid in which elementary excitations at low temperature are phonons with speed
e(p) u = lim , (16) T→0 p
such condition occurs when v < u. In our argument, we intentionally dismissed the role of momentum transfer by the population of thermal elementary phonons that always exist in the liquid and that is capable of carrying momentum from one point to another and thus provide a source of viscous friction. The momentum, which thermal elementary excitations may support, is finite and decrease with the temperature. It may be shown that such momentum is proportional to fluid velocity and is
P = Mexv. (17)
Whenever Mex < M the momentum capacity of thermal phonons becomes insufficient to decelerate the fluid, i.e., superfluidity is observed. In these conditions, one component of mass Mex acts like an ordinary fluid being subject to viscous forces and the remainder experiences zero viscous effect and forms the superfluid component. The amount of viscous components at a given temperature can be calculated in the microscopic theory of superfluidity, which we will describe below, and can be shown to fall as T4 as the temperature decreases. The microscopic theory of superfluidity had been suggested originally by N. Bogoli- ubov [60,63] and made use of three fundamental points, which are thought to be valid for low-temperature real fluid. The three fundamental assumptions of Bogoliubov’s treatment are that at low temperature a macroscopic number N0 of particles in the fluid occupies one single-particle state, for example, k = 0, the interaction between particles is essentially short range and its main effect consists in scattering particles in/out of the Bose condensate, thus making pairs of particles with momenta k and −k as the second largest population in the system. With these assumptions the original many-body Hamiltonian
0 † 1 † † H = e a ak + Vk −k a a ak ak δk +k k +k (18) ∑ k k 2V ∑ 1 3 k1 k2 3 4 1 2, 3 4 k k1,k2,k3,k4
can be transformed into
2 gN 1 h 0 † † † † i H = + ∑ (ek + ng)(ak ak + a−ka−k) + ng(ak a−k + aka−k) , (19) 2V 2 k6=0 Universe 2021, 7, 51 7 of 27
† where we also neglected “small” commutator [a0, a0] relative to N0. Here, g is the strength of effective short range interaction between the particles in the superfluid fluid and n = N/V is the particle density. It was further observed by N. Bogoliubov that Hamiltonian (19) can be exactly diago- nalized with the linear canonical transformation of the form (11) where
0 ! 2 2 1 ek + ng vk = uk − 1 = − 1 (20) 2 Ek
and q q 0 2 2 0 2 0 Ek = (ek + ng) − (ng) = (ek ) + 2ngek. (21) After BT, the Hamiltonian becomes
1 2 1 0 1 † † H = gn V − ∑ (ek + ng − Ek) + ∑ Ek(Ak Ak + A−k A−k). (22) 2 2 k6=0 2 k6=0
Equation (22) solves the original problem in its entirety yielding the energy spectrum of the elementary excitations, ground state energy and vacuum structure. In particular, all † elementary excitations correspond to coherent superpositions of particles ak and holes a−k in the Bose condensate with the wave vector k and energy Ek. In a long-wavelength limit (k → 0) the elementary excitations represent sound waves with the propagation speed
r ng s = . (23) M
For short wavelength, the spectrum is that of a free particle shifted upward by a constant gn arising from the interaction with the Bose condensate. From Equation (21), the thermal spectrum of the excitations can be easily derived and shown to be that of a Bose gas and all other thermodynamic properties of superfluid can be found [63]. In particular, one can easily show that the momentum, which can be carried by thermal excitations, at low temperatures is proportional to T4
Z d3 p P = − p · n(e − pv) ∼ vT4, (24) (2πh¯ )3
where n(e) is the Bose–Einstein distribution. In this brief example, the power of the application of CT can be vividly seen because an exact solution to a highly nontrivial problem is obtained with ease and full information about the system becomes available. Even when BT does not lead to a full diagonalization of the model Hamiltonian, the reduction gained by its use may be beneficial. In QCD, BT has been used continuously to describe nontrivial structure of a QCD vacuum, its superconducting properties and values of quark and glue condensations [64–67,71]. In a number of variational and field-theoretic works, it was suggested that the quantum configurations with quark pairs may have a lower free energy than a perturbative QCD vacuum and thus be energetically preferred leading to existence of nontrivial color condensation in a QCD vacuum. Such models had also been extended to describe properties of dressed constituent quarks and to derive from fundamental QCD hadron structure and mass spectrum [71]. In BT treatment, the QCD vacuum |Ωi is modeled as the BT vacuum annihilated by quasi-particle operators obtained from the original current-quark ladder operators with a rotation
† Aλ(k) = cos θkaλ(k) − λ sin θkbλ(k), † (25) Bλ(k) = cos θkbλ(k) + λ sin θkaλ(k). Universe 2021, 7, 51 8 of 27
Here, λ is the helicity of the quark. Such vacuum can be explicitly related to the perturbative vacuum by " # Z d3k |Ωi = exp − λ tan θ a† (k)b† (k) |0i, (26) ∑ ( )3 k λ λ λ 2π
where the gap angle θk is a free parameter used to minimize the energy of the trial QCD vacuum. The condition of the energy minimization typically results in a nonlinear integral gap equation, which needs to be solved in order for the explicit structure of the vacuum to become transparent. Such approach is able to describe the vacuum condensations in QCD as well as superconducting BCS features of the QCD vacuum. Operators Aλ(k) can be seen to describe the dressed quarks and may be further used to model dynamical mass generation of the constituent quarks in the Constituent Quark Model (CQM) and to produce a CQM-like description of the hadrons starting from the current quarks and fundamental QCD Hamiltonian [71]. The BT application to the QCD in 1+1 dimensions [72] has been exhibited for the mass gap of the quark [73–75] and the quark–antiquark bound-state equations [76–79]. Along this line, one employs the above QCD vacuum to construct meson states as produced on top of |Ωi by means of the meson creation operator
Z d3k Q† = ΨnJP(k)A† (k)B†(−k). (27) nJP ∑ ( )3 γδ λ δ γδ 2π
The application of a variational principle, or Tamm–Dancoff truncation, to
† |nJPi = QnJP|Ωi (28)
leads to a Schrodinger type equation, which can be solved for the spectrum and wave- functions of the mesons [71]. Further improvements of this approach, which rely on the QCD vacuum improved by an introduction of two and four quasi-particle correlations, is known in QCD as Random Phase Approximation (RPA) [71,80,81]. RPA, based on taking into account particle–particle and particle–hole correlations in the mean field, also applied in many other areas of quantum physics. In all of these applications, BT serves as a powerful tool to help gain nonperturbative knowledge about quantum-mechanical and field-theoretical properties of physical systems. BT, stemming from the classical apparatus of CTs, is more an exception rather than a rule given that the systematic translation of classical CTs into quantum framework experiences detrimental difficulties. Nonetheless, since its introduction in 1947, BT found a wide area of applications ranging from problems of condensed matter physics to strong interacting QCD as well as meson physics. It is repeatedly employed to describe nontrivial correlations in physical systems responsible for some dramatic physical behavior such as superfluidity, superconductivity, phase transitions, nontrivial vacuum condensations. In this article, we will now focus our attention on the application of BT that appeared in the field theory of flavor oscillations. In the upcoming sections, we will examine this development in greater detail.
3. General Theory of Quantum Field Mixing Quantum field mixing is a fascinating phenomenon, first observed in weak interac- tions, where the interacting states of a particle are dramatically different from the free- propagation states. As a result, the particle, say, produced in a weak decay evolves over time into a drastically different weak-interaction state with a very different weak decay signature. One usually thinks of this phenomenon in terms of weak-interaction (or flavor A, B, ... ) eigenstates and free-propagation (or energy a, b, ... ) eigenstates. The flavor state produced in weak interaction shall be represented as a superposition of energy-eigenstates, which then propagate independently from each other. Should weak decay happen once Universe 2021, 7, 51 9 of 27
more, however, the evolved superposition of energy-eigenstates should be represented again in terms of flavor-eigenstates to find appropriate weak-decay modes. In quantum mechanics, mixing of flavors is described by the interaction Hamiltonian 1
1 † † HI = ∑ ∑ mµν(aµ,kaν,k + aν,kaµ,k), (29) 2 k µ,ν=A,B,...
† where aA,B,...(aA,B,...) are creation (annihilation) operators for quantum flavor-states. The full Hamiltonian † H = ∑ ∑ eµ,kaµ,kaµ,k + HI (30) k µ=A,B,... can be straightforwardly diagonalized by introducing quantum energy states
† H = ∑ ∑ ei,kai,kai,k+const k i=a,b,... † (31) ai,k = ∑ Uiµaµ,k, µ=A,B,...
† where Uiµ is an appropriate unitary mixing matrix. This simple transformation allows one to immediately solve for time dynamics of flavor states in quantum mechanics and arrive to oscillation formulas for probability for a flavor state A to appear as a flavor state B after time t; |µ, ki = ∑ Uµi|i, ki i (32) 2 † −iei,kt 2 |hν, k|µ, k, ti| = | ∑ UµiUiνe | . i In the case of only two flavors, one recovers the famous Pontecorvo oscillation formu- las − |hA|A; ti|2 = 1 − sin2(2θ) sin2( Ea Eb t); 2 (33) 2 2 2 Ea−Eb |hB|A; ti| = sin (2θ) sin ( 2 t), where θ is a 2-flavor mixing angle. In quantum field theory, analogously, mixing is described with interaction Hamilto- nian density 1 † HI (φ(x)) = ∑ mµνφµ(x)φν(x) + h.c. (34) 2 µ,ν=A,B,... Full classical Hamiltonian can be similarly diagonalized with appropriately chosen linear transformation from flavor-fields φµ to mass-fields ϕi
† φµ → ϕi = ∑ Uiµφµ, (35) µ=A,B,...
0 H0(φ(x)) + HI (φ(x)) → H0(ϕ(x)). (36)
Here H0 is a free-theory Hamiltonian and the free-fields ϕi are given in terms of their Fourier transform as usual Z dk √ i ikx i † −ikx ϕi = ∑ ukσaikσ(t)e + vkσbikσ(t)e . (37) σ 2eik
1 We use the Latin indices i, j, k, ... and small Latin letters a, b, ... to label the mass-eigenstates and the Greek indices µ, ν, ρ, ... and capital Latin letters A, B, . . . to label the flavor-eigenstates. Universe 2021, 7, 51 10 of 27
−ie t −ie t aikσ(t) = e ik aikσ and bikσ(t) = e ik bikσ with the standard equal time commuta- i i tion/anticommutation relationships. In Equation (37), ukσ and vkσ are the free particle and antiparticle amplitudes, respectively, and σ is the helicity quantum number given by
i i i i (n · s)ukσ = σukσ ; (n · s)vkσ = σvkσ , (38)
where s is the spin operator and n = k/|k|. Differently from the quantum mechanics, in the quantum field theory the transforma- tion (35) does not immediately imply a specific form for the mixing relations between flavor and energy eigenstates. In fact, the intrinsic presence of antiparticle degrees of freedom in quantum field theory introduces a dramatic difference. In field theory any operators with the same conserved quantum numbers can mix. This means that in general in Equation (31) not only the flavor particle annihilation operators with momentum k and helicity σ will mix, but also the flavor antiparticle creation operators with momentum −k and helicity −σ may enter. Thus, the most general linear mixing relations in quantum field theory are