Θεώρημα Noether

Laws of physics


Emma Noether


Θεώρημα Noether Συμμετρία


Επιστημονικός Νόμος Επιστημονικοί Νόμοι
Μαθηματικό Θεώρημα Νόμοι Μαθηματικών
Φυσικός Νόμος Νόμοι Φυσικής
Νόμοι Χημείας
Νόμοι Γεωλογίας
Νόμοι Βιολογίας
Νόμοι Οικονομίας


Φυσικοί Γης Νόμοι Φυσικής Νόμοι Φυσικής Θεωρίες Φυσικής Πειράματα Φυσικής Παράδοξα ΦυσικήςΠροβλήματα Φυσικής

- Ένας Νόμος της Φυσικής.

- Ακριβέστερα, είναι ένας νόμος της Ρευστοδυναμικής

- Χρονολογία ανακάλυψης.


Πρότυπο:Physical Laws Η ονομασία "νόμος" σχετίζεται ετυμολογικά με το όνομα του φυσικού επιστήμονα "[[ ]]".


Για κάθε συνεχή συμμετρία που διέπει ένα ιδιαίτερο φυσικό φαινόμενο, (αναλλοίωτη η αντίστοιχη Lagrange), υπάρχει μία φυσική ποσότητα η οποία διατηρείται και δίνεται απο το γεννήτορα της αντίστοιχης ομάδας.


Το θεώρημα της Noether εκφράζει την ισοδυναμία που υπάρχει μεταξύ των νόμων διατήρησης και του αναλλοίωτου των φυσικών νόμων κάτω από ορισμένους μετασχηματισμούς (συνήθως αναφέρονται ως "συμμετρίες").

Για παράδειγμα:

Οι νόμοι διατήρησης επομένως είναι απόρροια αλλά και θεμέλιο της κατανόησης του Χωρόχρονου όπως τον ξέρουμε σήμερα και ισχύουν, τουλάχιστον μέχρι εκεί που μπορεί να φθάσει η ακρίβεια των πειραμάτων μας.


Noether's (first) theorem states that any differentiable symmetry of the action of a physical system has a corresponding conservation law. The theorem was proved by German mathematician Emmy Noether in 1915 and published in 1918.[1] The action of a physical system is the integral over time of a Lagrangian function (which may or may not be an integral over space of a Lagrangian density function), from which the system's behavior can be determined by the principle of least action.

Noether's theorem has become a fundamental tool of modern theoretical physics and the calculus of variations. A generalization of the seminal formulations on constants of motion in Lagrangian and Hamiltonian mechanics (1788 and 1833, respectively), it does not apply to systems that cannot be modeled with a Lagrangian; for example, dissipative systems with continuous symmetries need not have a corresponding conservation law.

For illustration, if a physical system behaves the same regardless of how it is oriented in space, its Lagrangian is rotationally symmetric; from this symmetry, Noether's theorem shows the angular momentum of the system must be conserved. The physical system itself need not be symmetric; a jagged asteroid tumbling in space conserves angular momentum despite its asymmetry – it is the laws of motion that are symmetric. As another example, if a physical experiment has the same outcome regardless of place or time (having the same outcome, say, somewhere in Asia on a Tuesday or in America on a Wednesday), then its Lagrangian is symmetric under continuous translations in space and time; by Noether's theorem, these symmetries account for the conservation laws of linear momentum and energy within this system, respectively. (These examples are just for illustration; in the first one, Noether's theorem added nothing new – the results were known to follow from Lagrange's equations and from Hamilton's equations.)

Noether's theorem is important, both because of the insight it gives into conservation laws, and also as a practical calculational tool. It allows researchers to determine the conserved quantities from the observed symmetries of a physical system. Conversely, it allows researchers to consider whole classes of hypothetical Lagrangians to describe a physical system. For illustration, suppose that a new field is discovered that conserves a quantity X. Using Noether's theorem, the types of Lagrangians that conserve X because of a continuous symmetry can be determined, and then their fitness judged by other criteria.

There are numerous different versions of Noether's theorem, with varying degrees of generality. The original version only applied to ordinary differential equations (particles) and not partial differential equations (fields). The original versions also assume that the Lagrangian only depends upon the first derivative, while later versions generalize the theorem to Lagrangians depending on the nth derivative. There is also a quantum version of this theorem, known as the Ward–Takahashi identity. Generalizations of Noether's theorem to superspaces also exist.

Informal statement of the theorem Edit

All fine technical points aside, Noether's theorem can be stated informally

If a system has a continuous symmetry property, then there are corresponding quantities whose values are conserved in time.[2]

A more sophisticated version of the theorem states that:

To every differentiable symmetry generated by local actions, there corresponds a conserved current.

The word "symmetry" in the above statement refers more precisely to the covariance of the form that a physical law takes with respect to a one-dimensional Lie group of transformations satisfying certain technical criteria. The conservation law of a physical quantity is usually expressed as a continuity equation.

The formal proof of the theorem uses only the condition of invariance to derive an expression for a current associated with a conserved physical quantity. The conserved quantity is called the Noether charge and the flow carrying that 'charge' is called the Noether current. The Noether current is defined up to a solenoidal vector field.

Historical contextEdit

A conservation law states that some quantity X describing a system remains constant throughout its motion; expressed mathematically, the rate of change of X (its derivative with respect to time) is zero:

$ \frac{dX}{dt} = 0 $

Such quantities are said to be conserved; they are often called constants of motion, although motion per se need not be involved, just evolution in time. For example, if the energy of a system is conserved, its energy is constant at all times, which imposes a constraint on the system's motion and may help to solve for it. Aside from the insight that such constants of motion give into the nature of a system, they are a useful calculational tool; for example, an approximate solution can be corrected by finding the nearest state that satisfies the necessary conservation laws.

The earliest constants of motion discovered were momentum and energy, which were proposed in the 17th century by René Descartes and Gottfried Leibniz on the basis of collision experiments, and refined by subsequent researchers. Isaac Newton was the first to enunciate the conservation of momentum in its modern form, and showed that it was a consequence of Newton's third law; interestingly, conservation of momentum still holds even in situations when Newton's third law is incorrect. Modern physics has revealed that the conservation laws of momentum and energy are only approximately true, but their modern refinements – the conservation of four-momentum in special relativity and the zero covariant divergence of the stress-energy tensor in general relativity – are rigorously true within the limits of those theories. The conservation of angular momentum, a generalization to rotating rigid bodies, likewise holds in modern physics. Another important conserved quantity, discovered in studies of the celestial mechanics of astronomical bodies, was the Laplace-Runge-Lenz vector.

In the late 18th and early 19th centuries, physicists developed more systematic methods for discovering conserved quantities. A major advance came in 1788 with the development of Lagrangian mechanics, which is related to the principle of least action. In this approach, the state of the system can be described by any type of generalized coordinates q; the laws of motion need not be expressed in a Cartesian coordinate system, as was customary in Newtonian mechanics. The action is defined as the time integral I of a function known as the Lagrangian L

$ I = \int L(\mathbf{q}, \dot{\mathbf{q}}, t) \, dt $

where the dot over q signifies the rate of change of the coordinates q

$ \dot{\mathbf{q}} = \frac{d\mathbf{q}}{dt}. $

Hamilton's principle states that the physical path q(t) – the one truly taken by the system – is a path for which infinitesimal variations in that path cause no change in I, at least up to first order. This principle results in the Euler–Lagrange equations

$ \frac{d}{dt} \left( \frac{\partial L}{\partial \dot{\mathbf{q}}} \right) = \frac{\partial L}{\partial \mathbf{q}}. $

Thus, if one of the coordinates, say qk, does not appear in the Lagrangian, the right-hand side of the equation is zero, and the left-hand side shows that

$ \frac{d}{dt} \left( \frac{\partial L}{\partial \dot{q}_k} \right) = \frac{dp_k}{dt} = 0 $

where the conserved momentum pk is defined as the left-hand quantity in parentheses. The absence of the coordinate qk from the Lagrangian implies that the Lagrangian is unaffected by changes or transformations of qk; the Lagrangian is invariant, and is said to exhibit a kind of symmetry. This is the seed idea from which Noether's theorem was born.

Several alternative methods for finding conserved quantities were developed in the 19th century, especially by William Rowan Hamilton. For example, he developed a theory of canonical transformations that allowed researchers to change coordinates so that coordinates disappeared from the Lagrangian, resulting in conserved quantities. Another approach and perhaps the most efficient for finding conserved quantities is the Hamilton-Jacobi equation.

Mathematical expressionEdit

The essence of Noether's theorem is the following: Imagine that the action I defined above is invariant under small perturbations (warpings) of the time variable t and the generalized coordinates q; (in a notation commonly used by physicists) we write

$ t \rightarrow t^{\prime} = t + \delta t $
$ \mathbf{q} \rightarrow \mathbf{q}^{\prime} = \mathbf{q} + \delta \mathbf{q} $

where the perturbations δt and δq are both small but variable. For generality, assume that there might be several such symmetry transformations of the action, say, N; we may use an index r = 1, 2, 3, …, N to keep track of them. Then a generic perturbation can be written as a linear sum of the individual types of perturbations

$ \delta t = \sum_r \epsilon_r T_r \! $
$ \delta \mathbf{q} = \sum_r \epsilon_r \mathbf{Q}_r. $

Using these definitions, Emmy Noether showed that the N quantities

$ \left(\frac{\partial L}{\partial \dot{\mathbf{q}}} \cdot \dot{\mathbf{q}} - L \right) T_r - \frac{\partial L}{\partial \dot{\mathbf{q}}} \cdot \mathbf{Q}_r $

are conserved, i.e., are constants of motion; this is a simple version of Noether's theorem.

Examples Edit

For illustration, consider a Lagrangian that does not depend on time, i.e., that is invariant (symmetric) under changes tt + δt, without any change in the coordinates q. In this case, N = 1, T = 1 and Q = 0; the corresponding conserved quantity is the total energy H[3]

$ H = \frac{\partial L}{\partial \dot{\mathbf{q}}} \cdot \dot{\mathbf{q}} - L. $

Similarly, consider a Lagrangian that does not depend on a coordinate qk, i.e., that is invariant (symmetric) under changes qkqk + δqk. In that case, N = 1, T = 0, and Qk = 1; the conserved quantity is the corresponding momentum pk[4]

$ p_k = \frac{\partial L}{\partial \dot{q_k}}. $

In special and general relativity, these apparently separate conservation laws are aspects of a single conservation law, that of the stress-energy tensor,[5] that is derived in the next section.

The conservation of the angular momentum L = r × p is slightly more complicated to derive, but analogous to its linear momentum counterpart.[6] It is assumed that the symmetry of the Lagrangian is rotational, i.e., that the Lagrangian does not depend on the absolute orientation of the physical system in space. For concreteness, assume that the Lagrangian does not change under small rotations of an angle δθ about an axis n; such a rotation transforms the Cartesian coordinates by the equation

$ \mathbf{r} \rightarrow \mathbf{r} + \delta\theta \mathbf{n} \times \mathbf{r}. $

Since time is not being transformed, T equals zero. Taking δθ as the ε parameter and the Cartesian coordinates r as the generalized coordinates q, the corresponding Q variables are given by

$ \mathbf{Q} = \mathbf{n} \times \mathbf{r}. $

Then Noether's theorem states that the following quantity is conserved

$ \frac{\partial L}{\partial \dot{\mathbf{q}}} \cdot \mathbf{Q}_{r} = \mathbf{p} \cdot \left( \mathbf{n} \times \mathbf{r} \right) = \mathbf{n} \cdot \left( \mathbf{r} \times \mathbf{p} \right) = \mathbf{n} \cdot \mathbf{L}. $

In other words, the component of the angular momentum L along the n axis is conserved. If n is arbitrary, i.e., if the system is insensitive to any rotation, then every component of L is conserved; in short, angular momentum is conserved.

Field-theory versionEdit

Although useful in its own right, the version of her theorem just given was a special case of the general version she derived in 1915. To give the flavor of the general theorem, a version of the Noether theorem for continuous fields in four-dimensional space-time is now given. Since field theory problems are more common in modern physics than mechanics problems, this field-theory version is the most commonly used version of Noether's theorem.

Let there be a set of differentiable fields φk defined over all space and time; for example, the temperature T(xt) would be representative of such a field, being a number defined at every place and time. The principle of least action can be applied to such fields, but the action is now an integral over space and time

$ I = \int L \left(\boldsymbol\phi, \partial_\mu{\boldsymbol\phi}, x^\mu \right) \, d^4 x $

(the theorem can actually be further generalized to the case where the Lagrangian depends on up to the nth derivative using jet bundles)

Let the action be invariant under certain transformations of the space-time coordinates xμ and the fields φk

$ x^{\mu} \rightarrow x^\mu + \delta x^\mu \! $
$ \boldsymbol \phi \rightarrow \boldsymbol \phi + \delta \boldsymbol \phi $

where the transformations can be indexed by r = 1, 2, 3, …, N

$ \delta x^\mu = \epsilon_r X^\mu_r \, $
$ \delta \boldsymbol\phi = \epsilon_r \boldsymbol\Psi_r $

For such systems, Noether's theorem states that there are N conserved current densities

$ j^\nu_r = - \left( \frac{\partial L}{\partial \boldsymbol\phi_{,\nu}} \right) \cdot \boldsymbol\Psi_r + \sum_{\sigma} \left[ \left( \frac{\partial L}{\partial \boldsymbol\phi_{,\nu}} \right) \cdot \boldsymbol\phi_{,\sigma} - L \delta^{\nu}_{\sigma} \right] X_{r}^{\sigma} $

In such cases, the conservation law is expressed in a four-dimensional way

$ \sum_\nu \frac{\partial j^\nu}{\partial x^\nu} = 0 $

which expresses the idea that the amount of a conserved quantity within a sphere cannot change unless some of it flows out of the sphere. For example, electric charge is conserved; the amount of charge within a sphere cannot change unless some of the charge leaves the sphere.

For illustration, consider a physical system of fields that behaves the same under translations in time and space, as considered above; in other words, the fields do not depend on the absolute position in space and time. In that case, N = 4, one for each dimension of space and time. Since only the positions in space-time are being warped, not the fields, the Ψ are all zero and the Xμν equal the Kronecker delta δμν, where we have used μ instead of r for the index. In that case, Noether's theorem corresponds to the conservation law for the stress-energy tensor Tμν[5]

$ T_\mu{}^\nu = \sum_{\sigma} \left[ \left( \frac{\partial L}{\partial \boldsymbol\phi_{,\nu}} \right) \cdot \boldsymbol\phi_{,\sigma} - L\,\delta^\nu_\sigma \right] \delta_\mu^\sigma = \left( \frac{\partial L}{\partial \boldsymbol\phi_{,\nu}} \right) \cdot \boldsymbol\phi_{,\mu} - L\,\delta_\mu^\nu $

The conservation of electric charge can be derived by considering transformations of the fields themselves.[7] In quantum mechanics, the probability amplitude ψ(x) of finding a particle at a point x is a complex field, because it ascribes a complex number to every point in space and time. The probability amplitude itself is physically unmeasurable; only the probability p = |ψ|2 is directly measureable. Therefore, the system is invariant under transformations of the ψ field and its complex conjugate field ψ* that leave |ψ|2 unchanged, such as

$ \psi \rightarrow e^{i\theta} \psi \ ,\ \psi^{*} \rightarrow e^{-i\theta} \psi^{*} $

In the limit when θ becomes infinitesimally small (δθ), it may be taken as the ε, and the Ψ are equal to iψ and −iψ*, respectively. A specific example is the Klein–Gordon equation, the relativistically correct version of the Schrödinger equation for spinless particles, which has the Lagrangian density

$ L = \psi_{,\nu} \psi^{*}_{,\mu} \eta^{\nu \mu} + m^2 \psi \psi^{*}. $

In this case, Noether's theorem states that the conserved current equals

$ j^{\nu} = i \left( \frac{\partial \psi}{\partial x^{\mu}} \psi^{*} - \frac{\partial \psi^{*}}{\partial x^{\mu}} \psi \right) \eta^{\nu \mu} $

which, when multiplied by the charge on that type of particle, equals the electric current density due to that type of particle. This transformation was first noted by Hermann Weyl and is one of the fundamental gauge symmetries of modern physics.

Derivations Edit

One independent variableEdit

Consider the simplest case, a system with one independent variable, time. Suppose the dependent variables $ \mathbf{q} $ are such that the action integral

$ I = \int_{t_1}^{t_2} L [\mathbf{q} [t], \dot{\mathbf{q}} [t], t] \, dt $

is invariant under brief infinitesimal variations in the dependent variables. In other words, they satisfy the Euler–Lagrange equations

$ \frac{d}{dt} \frac{\partial L}{\partial \dot{\mathbf{q}}} [t] = \frac{\partial L}{\partial \mathbf{q}} [t] . $

And suppose that the integral is invariant under a continuous symmetry. Mathematically such a symmetry is represented as a flow, $ \mathbf{\phi} $, which acts on the variables as follows

$ t \rightarrow t' = t + \epsilon T \! $
$ \mathbf{q} [t] \rightarrow \mathbf{q}' [t'] = \phi [\mathbf{q} [t], \epsilon] = \phi [\mathbf{q} [t' - \epsilon T], \epsilon] $

where $ \epsilon \! $ is a real variable indicating the amount of flow and T is a real constant (which could be zero) indicating how much the flow shifts time.

$ \dot {\mathbf{q}} [t] \rightarrow \dot{\mathbf{q}}' [t'] = \frac{d}{dt} \phi [\mathbf{q} [t], \epsilon] = \frac{\partial \phi}{\partial \mathbf{q}} [\mathbf{q} [t' - \epsilon T], \epsilon] \dot{\mathbf{q}} [t' - \epsilon T] . $

The action integral flows to

$ I' [\epsilon] = \int_{t_1 + \epsilon T}^{t_2 + \epsilon T} L [\mathbf{q}'[t'], \dot{\mathbf{q}}' [t'], t'] \, dt' $
$ I' [\epsilon] = \int_{t_1 + \epsilon T}^{t_2 + \epsilon T} L [\phi [\mathbf{q} [t' - \epsilon T], \epsilon], \frac{\partial \phi}{\partial \mathbf{q}} [\mathbf{q} [t' - \epsilon T], \epsilon] \dot{\mathbf{q}} [t' - \epsilon T], t'] \, dt' $

which may be regarded as a function of ε. Calculating the derivative at $ \epsilon = 0 $ and using the symmetry, we get

$ 0 = \frac{d I'}{d \epsilon} [0] = L [\mathbf{q} [t_2], \dot{\mathbf{q}} [t_2], t_2] T - L [\mathbf{q} [t_1], \dot{\mathbf{q}} [t_1], t_1] T \, + $
$ + \,\int_{t_1}^{t_2} \frac{\partial L}{\partial \mathbf{q}} \left( - \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} T + \frac{\partial \phi}{\partial \epsilon} \right) + \frac{\partial L}{\partial \dot{\mathbf{q}}} \left( - \frac{\partial^2 \phi}{(\partial \mathbf{q})^2} {\dot{\mathbf{q}}}^2 T + \frac{\partial^2 \phi}{\partial \epsilon \partial \mathbf{q}} \dot{\mathbf{q}} - \frac{\partial \phi}{\partial \mathbf{q}} \ddot{\mathbf{q}} T \right) \, dt $

Notice that the Euler–Lagrange equations imply

$ \frac{d}{dt} \left( \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} T \right) = $
$ = \left( \frac{d}{dt} \frac{\partial L}{\partial \dot{\mathbf{q}}} \right) \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} T + \frac{\partial L}{\partial \dot{\mathbf{q}}} \left( \frac{d}{dt} \frac{\partial \phi}{\partial \mathbf{q}} \right) \dot{\mathbf{q}} T + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \ddot{\mathbf{q}} \, T $
$ = \frac{\partial L}{\partial \mathbf{q}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} T + \frac{\partial L}{\partial \dot{\mathbf{q}}} \left( \frac{\partial^2 \phi}{(\partial \mathbf{q})^2} \dot{\mathbf{q}} \right) \dot{\mathbf{q}} T + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \ddot{\mathbf{q}} \, T. $

Substituting this into the previous equation, one gets

$ 0 = \frac{d I'}{d \epsilon} [0] = L [\mathbf{q} [t_2], \dot{\mathbf{q}} [t_2], t_2] T - L [\mathbf{q} [t_1], \dot{\mathbf{q}} [t_1], t_1] T - \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} [t_2] T + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} [t_1] T $
$ + \int_{t_1}^{t_2} \frac{\partial L}{\partial \mathbf{q}} \frac{\partial \phi}{\partial \epsilon} + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial^2 \phi}{\partial \epsilon \partial \mathbf{q}} \dot{\mathbf{q}} \, dt. $

Again using the Euler–Lagrange equations we get

$ \frac{d}{d t} \left( \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \epsilon} \right) = \left( \frac{d}{d t} \frac{\partial L}{\partial \dot{\mathbf{q}}} \right) \frac{\partial \phi}{\partial \epsilon} + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial^2 \phi}{\partial \epsilon \partial \mathbf{q}} \dot{\mathbf{q}} = \frac{\partial L}{\partial \mathbf{q}} \frac{\partial \phi}{\partial \epsilon} + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial^2 \phi}{\partial \epsilon \partial \mathbf{q}} \dot{\mathbf{q}}. $

Substituting this into the previous equation, one gets

$ 0 = L [\mathbf{q} [t_2], \dot{\mathbf{q}} [t_2], t_2] T - L [\mathbf{q} [t_1], \dot{\mathbf{q}} [t_1], t_1] T - \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} [t_2] T + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} [t_1] T + $
$ + \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \epsilon} [t_2] - \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \epsilon} [t_1]. $

From which one can see that

$ \left( \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \mathbf{q}} \dot{\mathbf{q}} - L \right) T - \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \epsilon} $

is a constant of the motion, i.e. a conserved quantity. Since $ \phi [\mathbf{q}, 0] = \mathbf{q} $, we get $ \frac{\partial \phi}{\partial \mathbf{q}} = 1 $ and so the conserved quantity simplifies to

$ \left( \frac{\partial L}{\partial \dot{\mathbf{q}}} \dot{\mathbf{q}} - L \right) T - \frac{\partial L}{\partial \dot{\mathbf{q}}} \frac{\partial \phi}{\partial \epsilon} . $

To avoid excessive complication of the formulas, this derivation assumed that the flow does not change as time passes. The same result can be obtained in the more general case.

Field-theoretic derivationEdit

Noether's theorem may also be derived for tensor fields φA where the index A ranges over the various components of the various tensor fields. These field quantities are functions defined over a four-dimensional space whose points are labeled by coordinates xμ where the index μ ranges over time (μ=0) and three spatial dimensions (μ=1,2,3). These four coordinates are the independent variables; and the values of the fields at each event are the dependent variables. Under an infinitesimal transformation, the variation in the coordinates is written

$ x^{\mu} \rightarrow \xi^{\mu} = x^{\mu} + \delta x^{\mu} \! $

whereas the transformation of the field variables is expressed as

$ {\phi}^A \rightarrow \alpha^A (\xi^{\mu}) = \phi^A (x^{\mu}) + \delta \phi^A (x^{\mu}) \,. $

By this definition, the field variations δφA result from two factors: intrinsic changes in the field themselves and changes in coordinates, since the transformed field αA depends on the transformed coordinates ξμ. To isolate the intrinsic changes, the field variation at a single point xμ may be defined

$ \alpha^A (x^{\mu}) = \phi^A (x^{\mu}) + \bar{\delta} \phi^A (x^{\mu}) \,. $

If the coordinates are changed, the boundary of the region of space-time over which the Lagrangian is being integrated also changes; the original boundary and its transformed version are denoted as Ω and Ω’, respectively.

Noether's theorem begins with the assumption that a specific transformation of the coordinates and field variables does not change the action, which is defined as the integral of the Lagrangian density over the given region of spacetime. Expressed mathematically, this assumption may be written as

$ \int_{\Omega^{\prime}} L \left( \alpha^A, {\alpha^A}_{,\nu}, \xi^{\mu} \right) d^{4}\xi - \int_{\Omega} L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) d^{4}x = 0 $

where the comma subscript indicates a partial derivative with respect to the coordinate(s) that follows the comma, e.g.

$ {\phi^A}_{,\sigma} = \frac{\partial \phi^A}{\partial x^{\sigma}} \,. $

Since ξ is a dummy variable of integration, and since the change in the boundary Ω is infinitesimal by assumption, the two integrals may be combined using the four-dimensional version of the divergence theorem into the following form

$ \int_{\Omega} \left\{ \left[ L \left( \alpha^A, {\alpha^A}_{,\nu}, x^{\mu} \right) - L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) \right] + \frac{\partial}{\partial x^{\sigma}} \left[ L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) \delta x^{\sigma} \right] \right\} d^{4}x = 0 \,. $

The difference in Lagrangians can be written to first-order in the infinitesimal variations as

$ \left[ L \left( \alpha^A, {\alpha^A}_{,\nu}, x^{\mu} \right) - L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) \right] = \frac{\partial L}{\partial \phi^A} \bar{\delta} \phi^A + \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \bar{\delta} {\phi^A}_{,\sigma} \,. $

However, because the variations are defined at the same point as described above, the variation and the derivative can be done in reverse order; they commute

$ \bar{\delta} {\phi^A}_{,\sigma} = \bar{\delta} \frac{\partial \phi^A}{\partial x^{\sigma}} = \frac{\partial}{\partial x^{\sigma}} \left( \bar{\delta} \phi^A \right) \,. $

Using the Euler–Lagrange field equations

$ \frac{\partial}{\partial x^{\sigma}} \left( \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \right) = \frac{\partial L}{\partial \phi^A} $

the difference in Lagrangians can be written neatly as

$ \left[ L \left( \alpha^A, {\alpha^A}_{,\nu}, x^{\mu} \right) - L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) \right] = \frac{\partial}{\partial x^{\sigma}} \left( \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \right) \bar{\delta} \phi^A + \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \bar{\delta} {\phi^A}_{,\sigma} = \frac{\partial}{\partial x^{\sigma}} \left( \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \bar{\delta} \phi^A \right) \,. $

Thus, the change in the action can be written as

$ \int_{\Omega} \frac{\partial}{\partial x^{\sigma}} \left\{ \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \bar{\delta} \phi^A + L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) \delta x^{\sigma} \right\} d^{4}x = 0 \,. $

Since this holds for any region Ω, the integrand must be zero

$ \frac{\partial}{\partial x^{\sigma}} \left\{ \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \bar{\delta} \phi^A + L \left( \phi^A, {\phi^A}_{,\nu}, x^{\mu} \right) \delta x^{\sigma} \right\} = 0 \,. $

For any combination of the various symmetry transformations, the perturbation can be written

$ \delta x^{\mu} = \epsilon X^{\mu} \! $
$ \delta \phi^A = \epsilon \Psi^A = \bar{\delta} \phi^A + \epsilon \mathcal{L}_X \phi^A $

where $ \mathcal{L}_X \phi^A $ is the Lie derivative of $ \phi^A \, $ in the $ X^\mu \, $ direction. When $ \phi^A \, $ is a scalar or $ {X^\mu}_{,\nu} = 0 \, $,

$ \mathcal{L}_X \phi^A = \frac{\partial \phi^A}{\partial x^{\mu}} X^{\mu} \,. $

These equations imply that the field variation taken at one point equals

$ \bar{\delta} \phi^A = \epsilon \Psi^A - \epsilon \mathcal{L}_X \phi^A \,. $

Differentiating the above divergence with respect to ε at ε=0 and changing the sign yields the conservation law

$ \frac{\partial }{\partial x^{\sigma}} j^{\sigma} = 0 $

where the conserved current equals

$ j^{\sigma} = \left[ \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \mathcal{L}_X \phi^A - L \, X^{\sigma} \right] - \left( \frac{\partial L}{\partial {\phi^A}_{,\sigma}} \right) \Psi^A \,. $

Manifold/fiber bundle derivationEdit

Suppose we have an n-dimensional oriented Riemannian manifold, M and a target manifold T. Let $ \mathcal{C} $ be the configuration space of smooth functions from M to T. (More generally, we can have smooth sections of a fiber bundle over M.)

Examples of this M in physics include:

  • In classical mechanics, in the Hamiltonian formulation, M is the one-dimensional manifold R, representing time and the target space is the cotangent bundle of space of generalized positions.
  • In field theory, M is the spacetime manifold and the target space is the set of values the fields can take at any given point. For example, if there are m real-valued scalar fields, φ1,…,φm, then the target manifold is Rm. If the field is a real vector field, then the target manifold is isomorphic to R3.

Now suppose there is a functional

$ \mathcal{S}:\mathcal{C}\rightarrow \mathbb{R}, $

called the action. (Note that it takes values into $ \mathbb{R} $, rather than $ \mathbb{C} $; this is for physical reasons, and doesn't really matter for this proof.)

To get to the usual version of Noether's theorem, we need additional restrictions on the action. We assume $ \mathcal{S}[\phi] $ is the integral over M of a function

$ \mathcal{L}(\phi,\partial_\mu\phi,x) $

called the Lagrangian density, depending on φ, its derivative and the position. In other words, for φ in $ \mathcal{C} $

$ \mathcal{S}[\phi]\,=\,\int_M \mathcal{L}[\phi(x),\partial_\mu\phi(x),x] \mathrm{d}^nx. $

Suppose we are given boundary conditions, i.e., a specification of the value of $ \phi\, $ at the boundary if M is compact, or some limit on $ \phi\, $ as x approaches ∞. Then the subspace of $ \mathcal{C} $ consisting of functions $ \phi\, $ such that all functional derivatives of $ \mathcal{S} $ at φ are zero, that is:

$ \frac{\delta \mathcal{S}[\phi]}{\delta \phi(x)}\approx 0 $

and that $ \phi\, $ satisfies the given boundary conditions, is the subspace of on shell solutions. (See principle of stationary action)

Now, suppose we have an infinitesimal transformation on $ \mathcal{C} $, generated by a functional derivation, Q such that

$ Q \left[ \int_N \mathcal{L} \, \mathrm{d}^n x \right] \approx \int_{\partial N} f^\mu [\phi(x),\partial\phi,\partial\partial\phi,\ldots] \mathrm{d}s_{\mu} $

for all compact submanifolds N or in other words,

$ Q[\mathcal{L}(x)]\approx\partial_\mu f^\mu(x) $

for all x, where we set

$ \mathcal{L}(x)=\mathcal{L}[\phi(x), \partial_\mu \phi(x),x].\ $

If this holds on shell and off shell, we say Q generates an off-shell symmetry. If this only holds on shell, we say Q generates an on-shell symmetry. Then, we say Q is a generator of a one parameter symmetry Lie group.

Now, for any N, because of the Euler–Lagrange theorem, on shell (and only on-shell), we have

$ Q\left[\int_N \mathcal{L} \, \mathrm{d}^nx \right] $ $ =\int_N \left[\frac{\partial\mathcal{L}}{\partial\phi}- \partial_\mu\frac{\partial\mathcal{L}}{\partial(\partial_\mu\phi)}\right]Q[\phi] \, \mathrm{d}^nx + \int_{\partial N} \frac{\partial\mathcal{L}}{\partial(\partial_\mu\phi)}Q[\phi] \, \mathrm{d}s_\mu $

$ \approx\int_{\partial N} f^\mu \, \mathrm{d}s_\mu . $

Since this is true for any N, we have

$ \partial_\mu\left[\frac{\partial\mathcal{L}}{\partial(\partial_\mu\phi)}Q[\phi]-f^\mu\right]\approx 0. $

But this is the continuity equation for the current $ J^\mu\,\! $ defined by:[8]

$ J^\mu\,=\,\frac{\partial\mathcal{L}}{\partial(\partial_\mu\phi)}Q[\phi]-f^\mu, $

which is called the Noether current associated with the symmetry. The continuity equation tells us that if we integrate this current over a space-like slice, we get a conserved quantity called the Noether charge (provided, of course, if M is noncompact, the currents fall off sufficiently fast at infinity).

Comments Edit

Noether's theorem is really a reflection of the relation between the boundary conditions and the variational principle. Assuming no boundary terms in the action, Noether's theorem implies that

$ \int_{\partial N} J^\mu \mathrm{d}s_\mu \approx 0. $

Noether's theorem is an on shell theorem. The quantum analog of Noether's theorem are the Ward–Takahashi identities.

Generalization to Lie algebras Edit

Suppose say we have two symmetry derivations Q1 and Q2. Then, [Q1Q2] is also a symmetry derivation. Let's see this explicitly. Let's say

$ Q_1[\mathcal{L}]\approx\partial_\mu f_1^\mu $


$ Q_2[\mathcal{L}]\approx\partial_\mu f_2^\mu $


$ [Q_1,Q_2][\mathcal{L}]=Q_1[Q_2[\mathcal{L}]]-Q_2[Q_1[\mathcal{L}]]\approx\partial_\mu f_{12}^\mu $

where f12=Q1[f2μ]-Q2[f1μ]. So,

$ j_{12}^\mu=\left(\frac{\partial}{\partial (\partial_\mu\phi)}\mathcal{L}\right)(Q_1[Q_2[\phi]]-Q_2[Q_1[\phi]])-f_{12}^\mu. $

This shows we can extend Noether's theorem to larger Lie algebras in a natural way.

Generalization of the proof Edit

This applies to any local symmetry derivation Q satisfying $ QS \approx 0 $, and also to more general local functional differentiable actions, including ones where the Lagrangian depends on higher derivatives of the fields. Let ε be any arbitrary smooth function of the spacetime (or time) manifold such that the closure of its support is disjoint from the boundary. ε is a test function. Then, because of the variational principle (which does not apply to the boundary, by the way), the derivation distribution q generated by q[ε][Φ(x)] = ε(x)Q[Φ(x)] satisfies $ q[\epsilon][S] \approx 0 $ for any ε, or more compactly, $ q(x)[S]\approx 0 $ for all x not on the boundary (but remember that q(x) is a shorthand for a derivation distribution, not a derivation parametrized by x in general). This is the generalization of Noether's theorem.

To see how the generalization related to the version given above, assume that the action is the spacetime integral of a Lagrangian that only depends on $ \phi\, $ and its first derivatives. Also, assume

$ Q[\mathcal{L}]\approx\partial_\mu f^\mu $


$ q[\epsilon][\mathcal{S}] = \int q[\epsilon][\mathcal{L}] \, \mathrm{d}^n x $
$ q[\epsilon][\mathcal{S}]= \int \left\{ \left(\frac{\partial}{\partial \phi}\mathcal{L}\right) \epsilon Q[\phi]+ \left[\frac{\partial}{\partial (\partial_\mu \phi)}\mathcal{L}\right]\partial_\mu(\epsilon Q[\phi]) \right\} \, \mathrm{d}^n x $
$ q[\epsilon][\mathcal{S}] = \int \left\{ \epsilon Q[\mathcal{L}] + \partial_{\mu}\epsilon \left[\frac{\partial}{\partial \left( \partial_{\mu} \phi\right)} \mathcal{L} \right] Q[\phi] \right\} \, \mathrm{d}^n x $
$ q[\epsilon][\mathcal{S}] \approx \int \epsilon \partial_\mu \Bigg\{f^\mu-\left[\frac{\partial}{\partial (\partial_\mu\phi)}\mathcal{L}\right]Q[\phi]\Bigg\} \, \mathrm{d}^n x $

for all ε.

More generally, if the Lagrangian depends on higher derivatives, then

$ \partial_\mu\left[f^\mu-\left[\frac{\partial}{\partial (\partial_\mu\phi)}\mathcal{L}\right]Q[\phi]-2\left[\frac{\partial}{\partial (\partial_\mu \partial_\nu \phi)}\mathcal{L}\right]\partial_\nu Q[\phi]+\partial_\nu\left[\left[\frac{\partial}{\partial (\partial_\mu \partial_\nu \phi)}\mathcal{L}\right] Q[\phi]\right]-\,\cdots\right]\approx 0. $

Examples Edit

Example 1: Conservation of energy Edit

Looking at the specific case of a Newtonian particle of mass m, coordinate x, moving under the influence of a potential V, coordinatized by time t. The action, S, is:

$ \mathcal{S}[x] = \int L[x(t),\dot{x}(t)] \, dt $
$ \mathcal{S}[x] = \int \left(\frac{m}{2}\sum_{i=1}^3\dot{x}_i^2-V(x(t))\right) \, dt. $

Consider the generator of time translations $ Q = \partial/\partial t $. In other words, $ Q[x(t)]=\dot{x}(t) $. Note that x has an explicit dependence on time, whilst V does not; consequently:

$ Q[L]=m \sum_i\dot{x}_i\ddot{x}_i-\sum_i\frac{\partial V(x)}{\partial x_i}\dot{x}_i = \frac{d}{dt}\left[\frac{m}{2}\sum_i\dot{x}_i^2-V(x)\right] $

so we can set

$ f=\frac{m}{2} \sum_i\dot{x}_i^2-V(x). $


$ j = \sum_{i=1}^3\frac{\partial L}{\partial \dot{x}_i}Q[x_i]-f $
$ j = m \sum_i\dot{x}_i^2 -\left[\frac{m}{2}\sum_i\dot{x}_i^2 -V(x)\right] \\ $
$ j = \frac{m}{2}\sum_i\dot{x}_i^2+V(x). $

The right hand side is the energy and Noether's theorem states that $ \dot{j}=0 $ (i.e. the principle of conservation of energy is a consequence of invariance under time translations).

More generally, if the Lagrangian does not depend explicitly on time, the quantity

$ \sum_{i=1}^3 \frac{\partial L}{\partial \dot{x}_i}\dot{x_i}-L $

(called the Hamiltonian) is conserved.

Example 2: Conservation of center of momentum Edit

Still considering 1-dimensional time, let

$ \mathcal{S}[\vec{x}] = \int \mathcal{L}[\vec{x}(t),\dot{\vec{x}}(t)] \, \mathrm{d}t $
$ \mathcal{S}[\vec{x}] = \int \left [\sum^N_{\alpha=1} \frac{m_\alpha}{2}(\dot{\vec{x}}_\alpha)^2 -\sum_{\alpha \beta} V_{\alpha\beta}(\vec{x}_\beta-\vec{x}_\alpha)\right] \, \mathrm{d}t $

i.e. N Newtonian particles where the potential only depends pairwise upon the relative displacement (α<β).

For $ \vec{Q} $, let's consider the generator of Galilean transformations (i.e. a change in the frame of reference). In other words,

$ Q_i[x^j_\alpha(t)]=t \delta^j_i. \, $

Note that

$ Q_i[\mathcal{L}] = \sum_\alpha m_\alpha \dot{x}_\alpha^i-\sum_{\alpha \beta}\partial_i V_{\alpha\beta}(\vec{x}_\beta-\vec{x}_\alpha)(t-t) $
$ Q_i[\mathcal{L}] = \sum_\alpha m_\alpha \dot{x}_\alpha^i. $

This has the form of $ \frac{\mathrm{d}}{\mathrm{d}t}\sum_\alpha m_\alpha x^i_\alpha $ so we can set

$ \vec{f}=\sum_\alpha m_\alpha \vec{x}_\alpha. $


$ \vec{j}=\sum_\alpha \left(\frac{\partial}{\partial \dot{\vec{x}}_\alpha}\mathcal{L}\right)\cdot\vec{Q}[\vec{x}_\alpha]-\vec{f} $
$ =\sum_\alpha (m_\alpha \dot{\vec{x}}_\alpha t-m_\alpha \vec{x}_\alpha) $
$ =\vec{P}t-M\vec{x}_{CM} $

where $ \vec{P} $ is the total momentum, M is the total mass and $ \vec{x}_{CM} $ is the center of mass. Noether's theorem states:

$ \dot{\vec{j}} = 0 \Rightarrow {\vec{P}}-M \dot{\vec{x}}_{CM} = 0. $

Example 3: Conformal transformation Edit

Both examples 1 and 2 are over a 1-dimensional manifold (time). An example involving spacetime is a conformal transformation of a massless real scalar field with a quartic potential in (3 + 1)-Minkowski spacetime.

$ \mathcal{S}[\phi]\, $ $ =\int \mathcal{L}[\phi (x),\partial_\mu \phi (x)] \, \mathrm{d}^4x $
$ =\int \left( \frac{1}{2}\partial^\mu \phi \partial_\mu \phi -\lambda \phi^4\right ) \, \mathrm{d}^4x $

For Q, consider the generator of a spacetime rescaling. In other words,

$ Q[\phi(x)]=x^\mu\partial_\mu \phi(x)+\phi(x). \! $

The second term on the right hand side is due to the "conformal weight" of φ. Note that

$ Q[\mathcal{L}]=\partial^\mu\phi\left(\partial_\mu\phi+x^\nu\partial_\mu\partial_\nu\phi+\partial_\mu\phi\right)-4\lambda\phi^3\left(x^\mu\partial_\mu\phi+\phi\right). $

This has the form of

$ \partial_\mu\left[\frac{1}{2}x^\mu\partial^\nu\phi\partial_\nu\phi-\lambda x^\mu\phi^4\right]=\partial_\mu\left(x^\mu\mathcal{L}\right) $

(where we have performed a change of dummy indices) so set

$ f^\mu=x^\mu\mathcal{L}.\, $


$ j^\mu=\left[\frac{\partial}{\partial (\partial_\mu\phi)}\mathcal{L}\right]Q[\phi]-f^\mu $
$ =\partial^\mu\phi\left(x^\nu\partial_\nu\phi+\phi\right)-x^\mu\left(\frac{1}{2}\partial^\nu\phi\partial_\nu\phi-\lambda\phi^4\right). $

Noether's theorem states that $ \partial_\mu j^\mu = 0 \! $ (as one may explicitly check by substituting the Euler–Lagrange equations into the left hand side).

(Aside: If one tries to find the Ward–Takahashi analog of this equation, one runs into a problem because of anomalies.)

Εφαρμογές Edit

Application of Noether's theorem allows physicists to gain powerful insights into any general theory in physics, by just analyzing the various transformations that would make the form of the laws involved invariant. For example:

In quantum field theory, the analog to Noether's theorem, the Ward–Takahashi identity, yields further conservation laws, such as the conservation of electric charge from the invariance with respect to a change in the phase factor of the complex field of the charged particle and the associated gauge of the electric potential and vector potential.

The Noether charge is also used in calculating the entropy of stationary black holes.[9]


  1. Noether E (1918). "Invariante Variationsprobleme". Nachr. D. König. Gesellsch. D. Wiss. Zu Göttingen, Math-phys. Klasse 1918: 235–257. doi:10.1080/00411457108231446. Bibcode1971TTSP....1..186N. 
  2. Thompson, W. J. (1994) Angular Momentum. Wiley, p.5.
  3. Lanczos, pp. 401–403.
  4. Lanczos, pp. 403–404.
  5. 5,0 5,1 Goldstein, pp. 592–593.
  6. Lanczos, pp. 404–405.
  7. Goldstein, pp.593–594.
  8. Michael E. Peskin, Daniel V. Schroeder (1995). An Introduction to Quantum Field Theory. Basic Books. σελ. 18. ISBN 0201503972. 
  9. Vivek Iyer; Wald (1995). "A comparison of Noether charge and Euclidean methods for Computing the Entropy of Stationary Black Holes". Physical Review D 52 (8): 4430–4439. doi:10.1103/PhysRevD.52.4430. Bibcode1995PhRvD..52.4430I. 

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Ikl Κίνδυνοι ΧρήσηςIkl

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