Renormalization

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In quantum field theory (QFT) and the statistical mechanics of fields, renormalization refers to a collection of techniques used to express physical calculations in terms of observable quantities that already include some field effects. Renormalization arose in quantum electrodynamics as a means of making sense of the infinite results of various calculations and extracting finite answers to properly posed physical questions. Initially viewed as a suspect, provisional procedure by most of its originators, renormalization eventually was embraced as an important tool in several fields of physics, as a result of work in effective field theory and the renormalization group.
Contents 
Prehistory: Selfinteractions in classical mechanics
Some of the problems and phenomena eventually addressed by renormalization actually appeared earlier in the classical electrodynamics of point particles in the 19th and early 20th century. When calculating the electromagnetic interactions of charged particles, one often ignores the backreaction of a particle's own field upon itself. It was realized early on that such a treatment is incomplete, and some theorists explored the intriguing idea that an electron's inertial mass could be entirely due to the backreaction. However, if the electron is assumed to be a point, the calculated value of this backreaction diverges, essentially because of the singularity at the origin in the inversesquare law. One potential solution was to assume that, say, the electron had a nonzero size, comparable to the number known as the classical electron radius, about 2.8 x 10^{15}m. Then, however, according to Henri Poincaré, the theory became inconsistent unless the electron possessed additional forces to hold it together internally against the repulsion of like charges. Today the hypothesis of a classical electron radius might be seen as an early attempt at regularization. Attempts to deal with the backreaction, such as the AbrahamLorentz theory, were never entirely internally consistent and exhibited bizarre phenomena such as anticausal "preacceleration", in which an electron would start moving shortly before a force was applied (Jackson 1998).
In classical field theory, therefore, the contribution of field interactions to a particle's physical properties was already problematic. Indeed, in some ways the trouble was worse than in QFT, since the shortdistance divergences involved were typically stronger than the ones encountered in quantum theories.
Divergences in quantum electrodynamics
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When developing quantum electrodynamics in the 1940s, Shin'ichiro Tomonaga, Julian Schwinger, Richard Feynman, and Freeman Dyson soon discovered that problems with divergent integrals abounded. One way of describing the divergences in QED is that they appear as the consequence of calculations involving Feynman diagrams with closed loops of virtual particles in them. One type of loop would be a situation in which a virtual electronpositron pair appear out of the vacuum, interact with various photons, and then annihilate. Another would be an electronphoton interaction as in Figure 2.
While virtual particles obey conservation of energy and momentum, they can possess combinations of energies and momenta not allowed by the classical laws of motion; physicists say that they are not on shell. Furthermore, whenever a loop appears, the particles involved in the loop are not individually constrained by the energies and momenta of incoming and outgoing particles, since a variation in, say, the energy of one particle in the loop can be balanced by an equal and opposite variation in the energy of another. Therefore, in order to calculate the contribution to a probability amplitude, one must integrate over all possible combinations of energy and momentum in the loop—and these integrals are often divergent, that is, they give infinite answers. The most theoretically troublesome divergences are the "ultraviolet" (UV) ones associated with large energies and momenta of the virtual particles in the loop, or, equivalently, very short wavelengths and high frequencies of the fields for which these particles are the quanta. These divergences are, therefore, fundamentally shortdistance, shorttime phenomena. (There are infrared divergences as well, but they are not as hard to interpret and are beyond the scope of this article.)
A loop divergence
The diagram in Figure 2 shows one of the several oneloop contributions to electronelectron scattering in QED. The electron on the left side of the diagram, represented by the solid line, starts out with fourmomentum <math>p^\mu<math> and ends up with fourmomentum <math>r^\mu<math>. It emits a virtual photon carrying <math>r^\mu  p^\mu<math> to transfer energy and momentum to the other electron. But in this diagram, before that happens, it emits another virtual photon carrying fourmomentum <math>q^\mu<math>, and it reabsorbs this one after emitting the other virtual photon. Energy and momentum conservation do not determine the fourmomentum <math>q^\mu<math> uniquely, so all possibilities contribute equally and we must integrate.
This diagram's amplitude ends up with, among other things, a factor from the loop of
 <math>
ie^3 \int {d^4 q \over (2\pi)^4} \gamma^\mu {i (\gamma^\alpha (rq)_\alpha + m) \over (rq)^2  m^2 + i \epsilon} \gamma^\rho {i (\gamma^\beta (pq)_\beta + m) \over (pq)^2  m^2 + i \epsilon} \gamma^\nu {i g_{\mu\nu} \over q^2 + i\epsilon } <math>
The various <math>\gamma^\mu<math> factors in this expression are gamma matrices as in the covariant formulation of the Dirac equation; they have to do with the spin of the electron. The important part for our purposes is the dependency on <math>q^\mu<math> of the three big factors in the integrand, which are from the propagators of the two electron lines and the photon line in the loop.
This has a piece with two powers of <math>q^\mu<math> on top that dominates at large values of <math>q^\mu<math> (Pokorski 1987, p. 122):
 <math>
e^3 \gamma^\mu \gamma^\alpha \gamma^\rho \gamma^\beta \gamma_\mu \int {d^4 q \over (2\pi)^4}{q_\alpha q_\beta \over (rq)^2 (pq)^2 q^2} <math>
This integral is divergent, and infinite unless we cut it off at finite energy and momentum in some way.
Similar loop divergences occur in other quantum field theories.
Renormalized and bare quantities
The solution was to realize that the quantities initially appearing in the theory's formulae (such as the Lagrangian), representing such things as the electron's electric charge and mass, as well as the normalizations of the quantum fields themselves, did not actually correspond to the physical constants measured in the laboratory. As written, they were bare quantities that did not take into account the contribution of virtualparticle loop effects to the physical constants themselves. Among other things, these effects would include the quantum counterpart of the electromagnetic backreaction that so vexed classical theorists of electromagnetism. In general, these effects would be just as divergent as the amplitudes under study in the first place; so finite measured quantities would in general imply divergent bare quantities.
In order to make contact with reality, then, the formulae would have to be rewritten in terms of measurable, renormalized quantities. The charge of the electron, say, would be defined in terms of a quantity measured at a specific kinematic renormalization point or subtraction point (which will generally have a characteristic energy, called the renormalization scale or simply the energy scale). The parts of the Lagrangian left over, involving the remaining portions of the bare quantities, could then be reinterpreted as counterterms, involved in divergent diagrams exactly canceling out the troublesome divergences for other diagrams.
Renormalization in QED
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For example, in the Lagrangian of QED
 <math>\mathcal{L}=\bar\psi_B\left[i\gamma_\mu (\partial^\mu + ie_BA_B^\mu)m_B\right]\psi_B \frac{1}{4}F_{B\mu\nu}F_B^{\mu\nu}<math>
the fields and coupling constant are really bare quantities, hence the subscript <math>B<math> above. Conventionally the bare quantities are written so that the corresponding Lagrangian terms are multiples of the renormalized ones:
 <math>\left(\bar\psi m \psi\right)_B = Z_0 \, \bar\psi m \psi<math>
 <math>\left(\bar\psi (\partial^\mu + ieA^\mu)\psi\right)_B = Z_1 \, \bar\psi(\partial^\mu + ieA^\mu)\psi<math>
 <math>\left(F_{\mu\nu}F^{\mu\nu}\right)_B = Z_3\, F_{\mu\nu}F^{\mu\nu}<math>.
(Gauge invariance, via a WardTakahashi identity, turns out to imply that we can renormalize the two terms of the covariant derivative piece <math>\bar \psi (\partial + ieA) \psi<math> together (Pokorski 1987, p. 115), which is what happened to <math>Z_2<math>; it is the same as <math>Z_1<math>.)
A term in this Lagrangian, for example, the electronphoton interaction pictured in Figure 1, can then be written
 <math>\mathcal{L}_I = e \bar\psi \gamma_\mu A^\mu \psi \,  \, (Z_1  1) e \bar\psi \gamma_\mu A^\mu \psi<math>
The physical constant <math>e<math>, the electron's charge, can then be defined in terms of some specific experiment; we set the renormalization scale equal to the energy characteristic of this experiment, and the first term gives the interaction we see in the laboratory (up to small, finite corrections from loop diagrams, providing such exotica as the highorder corrections to the magnetic moment). The rest is the counterterm. If we are lucky, the divergent parts of loop diagrams can all be decomposed into pieces with three or fewer legs, with an algebraic form that can be canceled out by the second term (or by the similar counterterms that come from <math>Z_0<math> and <math>Z_3<math>). In QED, we are lucky: the theory is renormalizable (see below for more on this).
The diagram with the <math>Z_1<math> counterterm's interaction vertex placed as in Figure 3 cancels out the divergence from the loop in Figure 2.
The splitting of the "bare terms" into the original terms and counterterms came before the renormalization group insights due to Kenneth Wilson. According to the renormalization group insights, this splitting is unnatural and unphysical.
Running constants
To minimize the contribution of loop diagrams to a given calculation (and therefore make it easier to extract results), one chooses a renormalization point close to the energies and momenta actually exchanged in the interaction. However, the renormalization point is not itself a physical quantity: the physical predictions of the theory, calculated to all orders, should in principle be independent of the choice of renormalization point, as long as it is within the domain of application of the theory. Changes in renormalization scale will simply affect how much of a result comes from Feynman diagrams without loops, and how much comes from the leftover finite parts of loop diagrams. One can exploit this fact to calculate the effective variation of physical constants with changes in scale, also known as the renormalization group.
Colloquially, particle physicists often speak of certain physical constants as varying with the energy of an interaction, though in fact it is the renormalization scale that is the independent quantity. This running does, however, provide a convenient means of describing changes in the behavior of a field theory under changes in the energies involved in an interaction. For example, since the coupling constant in quantum chromodynamics becomes small at large energy scales, the theory behaves more like a free theory as the energy exchanged in an interaction becomes large, a phenomenon known as asymptotic freedom. Choosing an increasing energy scale and using the renormalization group makes this clear from simple Feynman diagrams; were this not done, the prediction would be the same, but would arise from complicated highorder cancellations.
Regularization
Since the quantity <math>\infty  \infty<math> is illdefined, in order to make this notion of canceling divergences precise, the divergences first have to be tamed mathematically using the theory of limits, in a process known as regularization.
An essentially arbitrary modification to the loop integrands, or regulator, can make them drop off faster at high energies and momenta, in such a manner that the integrals converge. A regulator has a characteristic energy scale known as the cutoff; taking this cutoff to infinity (or, equivalently, the corresponding length/time scale to zero) recovers the original integrals.
With the regulator in place, and a finite value for the cutoff, divergent terms in the integrals then turn into finite but cutoffdependent terms. After canceling out these terms with the contributions from cutoffdependent counterterms, the cutoff is taken to infinity and finite physical results recovered. If physics on scales we can measure is independent of what happens at the very shortest distance and time scales, then it should be possible to get cutoffindependent results for calculations.
Many different types of regulator are used in quantum field theory calculations, each with its advantages and disadvantages. One of the most popular in modern use is dimensional regularization, invented by Gerardus 't Hooft and Martinus J. G. Veltman, which tames the integrals by carrying them into a space with a fictitious fractional number of dimensions. Another is PauliVillars regularization, which adds fictitious particles to the theory with very large masses, such that loop integrands involving the massive particles cancel out the existing loops at large momenta.
Attitudes and interpretation
The early formulators of QED and other quantum field theories were, as a rule, dissatisfied with this state of affairs. It seemed illegitimate to do something tantamount to subtracting infinities from infinities to get finite answers. Richard Feynman famously called renormalization "a dippy process" (Feynman 1988) and suggested that it betrayed some deep lack of human understanding about quantum fields. The general unease was almost universal in texts up through the 1970s and 1980s, and has continued to some extent to the present day.
Beginning in the 1970s, however, inspired by work on the renormalization group and effective field theory, attitudes began to change, especially among younger theorists. Kenneth G. Wilson and others demonstrated that the renormalization group is useful in statistical field theory applied to condensed matter physics, where it provides important insights into the behavior of phase transitions. In condensed matter physics, a real shortdistance regulator exists: matter ceases to be continuous on the scale of atoms. Shortdistance divergences in condensed matter physics do not present a philosophical problem, since the field theory is only an effective, smoothedout representation of the behavior of matter anyway; there are no infinities since the cutoff is actually always finite, and it makes perfect sense that the bare quantities are cutoffdependent.
If we do not believe that QFT as we know it holds, say, all the way down past the Planck length (where it might yield to string theory or something stranger), then there may be no real problem with shortdistance divergences in particle physics either; all field theories could simply be effective field theories. In a sense, this approach echoes the older attitude that the divergences in QFT speak of human ignorance about the workings of nature, but also acknowledges that this ignorance can be quantified and that the resulting effective theories remain useful.
In QFT, the value of a physical constant, in general, depends on the scale that one chooses as the renormalization point, and it becomes very interesting to examine the renormalization group running of physical constants under changes in the energy scale. The coupling constants in the Standard Model of particle physics vary in different ways with increasing energy scale: the coupling of quantum chromodynamics and the weak isospin coupling of the electroweak force tend to decrease, and the weak hypercharge coupling of the electroweak force tends to increase. At the colossal energy scale of 10^{15} GeV (far beyond the reach of our civilization's particle accelerators), they all become approximately the same size (Grotz and Klapdor 1990, p. 254), a major motivation for speculations about grand unified theory. Instead of a worrisome problem, renormalization has become an important theoretical tool for studying the behavior of field theories in different regimes.
Renormalizability
Related to this philosophical sea change is the notion of renormalizability. Not all theories lend themselves to renormalization in the manner described above, with a finite supply of counterterms and all quantities becoming cutoffindependent at the end of the calculation. If the Lagrangian contains combinations of field operators of excessively high dimension in energy units, the counterterms required to cancel all divergences proliferate to infinite number, and, at first glance, the theory would seem to gain an infinite number of free parameters and therefore lose all predictive power, becoming scientifically worthless. Such theories are called nonrenormalizable.
The Standard Model of particle physics contains only renormalizable operators, but the interactions of general relativity become nonrenormalizable operators if one attempts to construct a field theory of quantum gravity in the most straightforward manner, suggesting that perturbation theory is useless in application to quantum gravity.
However, in an effective field theory, "renormalizability" is, strictly speaking, a misnomer. In a nonrenormalizable effective field theory, terms in the Lagrangian do multiply to infinity, but have coefficients suppressed by evermoreextreme inverse powers of the energy cutoff. If the cutoff is a real, physical quantity—if, that is, the theory is only an effective description of physics up to some maximum energy or minimum distance scale—then these extra terms could represent real physical interactions. Assuming that the dimensionless constants in the theory do not get too large, one can group calculations by inverse powers of the cutoff, and extract approximate predictions to finite order in the cutoff that still have a finite number of free parameters. It can even be useful to renormalize these "nonrenormalizable" interactions.
Nonrenormalizable interactions in effective field theories rapidly become weaker as the energy scale becomes much smaller than the cutoff. The classic example is the Fermi theory of the weak nuclear force, a nonrenormalizable effective theory whose cutoff is comparable to the mass of the W particle. This fact may also provide a possible explanation for why almost all of the particle interactions we see are describable by renormalizable theories. It may be that any others that may exist at the GUT or Planck scale simply become too weak to detect in the realm we can observe, with one exception: gravity, whose exceedingly weak interaction is magnified by the presence of the enormous masses of stars and planets.
References
 Feynman, Richard (1988). QED: The Strange Theory of Light and Matter. Princeton: Princeton University Press. ISBN 0691024170
 Griffiths, David (1989). Introduction to Elementary Particles. New York: Wiley. ISBN 0471603864
 Grotz, K., and Klapdor, H. V (1990). The Weak Interaction in Nuclear, Particle and Astrophysics. Wilson, S. S., trans. Bristol: Adam Hilger. ISBN 0852743130
 Itzykson, Claude, and Zuber, JeanBernard (1980). Quantum Field Theory. New York: McGrawHill. ISBN 0070320713
 Jackson, J. D. (1998). Classical Electrodynamics. New York: Wiley. ISBN 047130932X
 Pokorski, Stefan (1987). Gauge Field Theories. Cambridge: Cambridge University Press. ISBN 0521368464
External links
 Baez, John (1998). Renormalization Made Easy (http://math.ucr.edu/home/baez/renormalization.html). Retrieved October 17, 2004. A qualitative introduction to the subject.
 Blechman, Andrew E. (2002). Renormalization: Our Greatly Misunderstood Friend (http://www.pha.jhu.edu/~blechman/papers/renormalization/). Retrieved October 17, 2004. Summary of a lecture; has more information about specific regularization and divergencesubtraction schemes.
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