text:
FEATURED AUTHOR: JULIE ANNE LINDSEY
It's been three months since Jake Archer rolled into town, accused Mia Connors of murder and stole her heart. She hasn't heard from him since. So when a man collapses at the fall Renaissance Faire, she's surprised to see the US marshal arrive on scene. And shocked when he points the finger at her—again. Mia would sooner be able to resurrect the poor fellow than poison him.
Jake Archer's career has been rising fast, but it's about to come to a crashing halt. The Ren-Faire victim was in protective custody — Jake's custody — and they were painfully close to nabbing a major crime boss. If Jake doesn't solve the murder soon, he'll be fetching donuts instead of protecting his nation. A difficult enough task without the alluring Mia Connors in the way.
Working with Jake to catch the killer might push Mia into crazy-cat-lady territory. But with a murderer on the loose — and Mia's reputation on the line — they'll have to work fast to find the killer before the killer finds them.
INTERVIEW WITH JULIE ANNE LINDSEY
Julie, do you write every day?
Yes! I'm a dedicated outliner, and I write one chapter every day, reread it for cleanup, then send it to my critique partner. By following this process, I finish writing my novels in under a month. (Of course, that doesn't count the time spent prepping my outline, researching the project or reworking the novel after my beta readers have given feedback on the completed project). I'm a firm believer in writing every day. Writing daily builds momentum, finishes projects and give agents and editors a reliable time frame for completion.
How often do you read?
I read daily, though sometimes it isn't more than a few pages. Reading is so important. I read for research, for entertainment, and to better understand the industry. Reading teaches me new ways to say old things, gives me a little reprieve from the writing, and shows me what editors are buying, what readers are reading and where to aim my next project.
What books do you currently have published?
Currently, I have two cozy mystery series with Carina Press.
The Patience Price Mysteries:
• Murder by the Seaside
• Murder Comes Ashore
• Murder in Real Time
The Geek Girl Mysteries
• A Geek Girl's Guide to Murder
• A Geek Girl's Guide to Arsenic
• A Geek Girl's Guide to Justice (arrives this fall)
And several books for young adults:
• Deceived
• Goddess
• In Place of Never
• What She Wanted (arrives in July)
Is writing your dream job?
Absolutely. I love writing. I can't imagine life without it.
For what would you like to be remembered?
I'd like to be remembered as the author who made people smile.
The wife and mother who put God and her family first.
The friend who never lets you down.
What do you love about where you live?
I live in rural Ohio, and I love everything about it. There's little crime, beautiful seasons, a thriving Amish community, lots of open land and livestock. The library is nicely stocked and funded. The pace is slow. The neighbors are friendly, and the schools are fantastic. There's no other place I'd want to raise my family.
What's your favorite fast food?
Chipotle veggie bowl with fajitas, black beans, tomato salsa and guac!
What's your favorite beverage?
Coffee. I live for the coffee.
Where is your favorite place to visit?
New Orleans!
I visited the city last summer when one of my books was nominated for a mystery award, and I fell hopelessly in love. I extended the trip and spent the extra time learning everything I could about local history and culture. Inspiration was everywhere. I started outlining a new mystery on the flight home. That book is called Cat Got Your Diamonds. It's now finished! Sold! And in the fall lineup with Crooked Lane Books!
What's your least favorite chore? Why?
I loathe laundry. There are five in my family, and the laundry never ever ends. There's not even hope for an end. There will never be a feeling of accomplishment when I reach the bottom of the hamper. I don't think my hamper has one. Dirty clothes, towels, swim gear (all three kids are competitive swimmers), ski gear, etc just keeps coming. It's depressing to think about.
What's the hardest thing you've ever had to write?
Last summer, I had to write a novel about a teen whose grandfather had a massive heart attack. The novel was sold months earlier on proposal and the due date was fast approaching. Unfortunately, my life was in upheaval. We lost my father-in-law to a massive heart attack while I was writing the manuscript. My husband, children and heart were in utter turmoil. It killed me to write/relive our loss. The wound was too fresh. I took days off writing to support my family and mental health, attend memorials and the funeral, but I had to finish the manuscript. So, every day that I sat down to write, I'd lose my mind. My emotions would take over, tears would blur the screen and I'd end up in bed. That book is called What She Wanted, and it releases this July, a year after our devastating loss. I'm not sure I'll even be able to promote or talk about it when it arrives. A year later is still too soon.
What is the wallpaper on your computer's desktop?
My wallpaper is a summer pic of my three kids playing in the rain, fists raised skyward, mouths open. They went out to play, but didn't come in when the thunder started, so I went to call for them inside. I found them screaming, "By the power of Asgard!" (My family loves Marvel, especially Thor).
I'm writing book two in The Kitty Couture Mysteries! It's a fun new series set in the famed New Orleans Garden District with lots of pets, fun and flavor. My heroine, Lacy Marie Crocker, runs Furry Godmother, a pet boutique and bakery on Magazine Street. She's drawn into investigations that impact her close knit community of wealthy eccentrics and devoted pet lovers. Book one, Cat Got Your Diamonds, releases November 7, 2016!
Julie Anne Lindsey is a multi-genre author who writes the stories that keep her up at night. She's a self-proclaimed nerd with a penchant for words and proclivity for fun. Julie lives in rural Ohio with her husband and three small children. Today, she hopes to make someone smile. One day she plans to change the world. Julie is a member of the International Thriller Writers (ITW) and Sisters in Crime (SinC). She is represented by Jill Marsal of Marsal Lyons Literary Agency. Julie also writes as Julie Chase.
Connect with Julie Anne:
Labels: A Geek Girl's Guide to Arseni, cozy mystery, geek, Julie Anne Lindsey


Freehand filigree disc in random pattern with adjustable chain.
With a rich texture, this versatile necklace looks great worn short or long, especially accompanying a black dress.
Plain sterling silver, or with 18ct yellow gold vermeil.
Pendant size 3.5cm diameter. Chain length 45cm.


text:
Fame Doesn't Buy Happiness for Kurt in AGE OF X-MAN: AMAZING NIGHTCRAWLER #2
By Peyton Hinckle
AGE OF X-MAN: THE AMAZING NIGHTCRAWLER #2 by Seanan McGuire, Juan Frigeri, and Shane Davis
AGE OF X-MAN: AMAZING NIGHTCRAWLER #2 is successful in that it's entertaining. The characters and the plotlines are intriguing but the execution is where this issue falls apart. Frigeri and Davis struggle to use art to tell a cohesive story. With so many complex plotlines going on, the art has to be strong enough to hold it all and, for AGE OF X-MAN: AMAZING NIGHTCRAWLER #2, it isn't. Still, that doesn't mean this issue isn't worth a read. If you're a Nightcrawler fan, this one is still a must-buy.
User Rating 0 Be the first one !
Kurt's life takes a downward spiral in this week's AGE OF X-MAN: THE AMAZING NIGHTCRAWLER #2! While not quite as strong as the first issue, the second installment in Nightcrawler's spotlight miniseries does help develop some interesting storylines that could become something bigger down the line.
From the underground "love club" to the introduction of "Mastermind Studios," writer Seanan McGuire continues to broaden the world Nate Grey created with varying degrees of success. Artists Juan Frigeri and Shane Davis, unfortunately, hurt the issue more than help it. Partly why I can't give AGE OF X-MAN: THE AMAZING NIGHTCRAWLER #2 a better score is because of the art. Until that changes, this miniseries won't be half as memorable as it could be.
AGE OF X-MAN: AMAZING NIGHTCRAWLER #2 Review / Image courtesy of Marvel Entertainment
Kurt Doesn't Want the Love
This issue covers two main storylines: Kurt's romance (or lack thereof) with his co-star Meggan and the X-Men's film studio's feud with rival company Mastermind Studios. In AGE OF X-MAN: THE AMAZING NIGHTCRAWLER #1, Kurt and Meggan have sex which, as we all know by now, is a big no-no in Nate's world. In AGE OF X-MAN: THE AMAZING NIGHTCRAWLER #2, Kurt has to make a big decision: should he continue seeing Meggan romantically and face persecution or should he be a "good" X-Man and follow the rules? Ultimately, he wisely chooses the latter and leaves Meggan.
On the business side of things, Kurt's facing a barrage of threats from Regan Wyngarde a.k.a. Lady Mastermind. She wants her semi-corrupt film studio to team up with the X-Men's Studio X. Not surprisingly, Kurt isn't interested in a merger and bluntly turns her down.
Lady Mastermind Becomes a Business Woman
We see a lot of new characters in this issue, from Lady Mastermind and Surge to Julian Keller and Phoebe Cuckoo. Out of all of them, Reygan gets the most panel-time, since she owns the competing Mastermind Studios. I'm not the hugest fan of Reygan in the regular Earth-616 comics, but in AGE OF X-MAN: AMAZING NIGHTCRAWLER #2, McGuire knows how to handle her character.
Reygan's position as a ruthless, business-minded mastermind (pun intended) makes perfect sense. She's a great example of how to keep a villain a villain in a different reality without just adding loads and loads of physical violence. She's a bad guy, but in a different, more unique way. My only complaint is her outfit, which is both sexually gratuitous and out-of-character for her role in the issue. I needed a strong power-suit and instead, I got the skintight latex bodysuit she always wears.
Meggan Does Want the Love
I really enjoyed Lady Mastermind's characterization but her storyline (the potential merger between the two dueling studios) fell a little flat. For the other storyline, it turned out to be the opposite. Kurt and Meggan's budding relationship is an intriguing thread but Meggan's characterization feels strangely distant. We don't see a ton of her in the issue and when we do, her personality is all over the place. She should be Kurt's "co-star" instead of a lackluster supporting character.
With that being said, the evolution of Kurt and Meggan's relationship in AGE OF X-MAN: AMAZING NIGHTCRAWLER #2 is entertaining. The secret "love club" that Meggan invites Kurt to is a welcome addition to the ever-expanding Age of X-Man world. Little things like that help to make this reality all the more real (or as real as a comic book alternate reality can be).
Sexualization and Distortion in AGE OF X-MAN: AMAZING NIGHTCRAWLER #2
Drawing the interior art for comics is incredibly difficult. I don't doubt that artist Juan Frigeri is beyond talented. However, for this issue (and, seemingly, for this miniseries) his art just isn't functioning well with the storyline. Frigeri really struggles to draw female characters' facial expressions and figures realistically.
In most panels, Meggan's features are so fine and her eyes so small that she looks strangely distorted. Her body's proportions are also abnormally dramatic — even for a comic book. Most of the other female characters face a similar fate, but the male characters don't. Nightcrawler looks normal in the majority of his scenes. Although Kurt's the star, there are too many female characters in AGE OF X-MAN: AMAZING NIGHTCRAWLER #2 to ignore them.
Colorist Shane Davis does a good job delivering a fairly predictable palette. A more complex palette may have been able to elevate the pencils slightly, but, at the end of the day, the lines have to be strong for the colors to be strong.
The Age of X-Man Rages On
With each Age of X-Man issue, we get a little more information about this larger world. AGE OF X-MAN: AMAZING NIGHTCRAWLER #2 does a great job exploring that aspect of alternate reality building, but it fails in other categories, including art and characterization. If you love Nightcrawler, you probably already subscribed to this series. But, if you're more of a casual, overall X-Men fan, you don't necessarily need to buy this issue. Age of X-Man is a huge event. Not everyone can buy every single issue. Stick to AGE OF X-MAN: THE MARVELOUS X-MEN, where the larger events are centered.
TagsX-MenNightcrawlerSeanan McGuireage of x-manjuan frigeriAGE OF X-MAN: THE AMAZING NIGHTCRAWLER #2meggan
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redpajama_set_name: RedPajamaCommonCrawl


The Italian American community of Cleveland lost a leader, an intellectual, a gentleman, and a friend this past October 2018. Alfonso D'Emilia was 90-years-old when he passed away after a short illness. He was the beloved husband for 63 years of Mary Jane Volzer whom he met on a blind date to see the New York Opera Co. perform in Cleveland. Together they had 7 daughters: Mary, Florence, Patricia, Catherine, Ruth, Joni, and Jacki.
Alfonso was born in Colle Sanita, Benevento, Italy. Even though born in Italy, Alfonso was an American citizen because his father had immigrated to the U.S. in 1915 and naturalized. Alfonso immigrated to the U.S. in 1948 at the age of 20. He initially came to Cleveland to join his father, but soon left for New York to take a position as translator for IlProgresso Italo-Americano.
Alfonso came back to Cleveland a few years later to become an editor for the newspaper L'Araldo. Several years after that, he became a member of the staff of The Plain Dealer and had a successful and distinguished 44-year career with the paper as a book reviewer and national representative. Besides his work with the newspapers, Alfonso was a community leader very active with other endeavors.
As a historian and gifted author, Alfonso presented and lectured to local, national and international audiences. In 1968, he gave a series of lectures on Italian cultural heritage at Cuyahoga Community College and John Carroll University. In 1975, Alfonso participated at the International Congress for Studies on Christopher Columbus at the University of Genoa, Italy. His work and reputation earned him a seat on the National Executive Committee of the American Italian Historical Association.
Service on several boards and committees showed Alfonso's strength as a leader and advisor. As a past president of the Cleveland Cultural Gardens Federation, he led the ethnic community of Cleveland in caring for one of its most important cultural assets. He served on the board of the International Services Center, the Cultural Heritage Advisory Committee of Cuyahoga Community College and on both the Cleveland and Ohio Bicentennial Advisory Committees. Alfonso was a dedicated member of Il Cenacolo Italiano di Cleveland since the 1960s, serving in many leadership roles and frequently as a lecturer.
Recognitions for Alfonso's outstanding contributions and character include: being bestowed Cavaliere della Repubblica d'Italia (Knight of the Republic of Italy) in 1971, a proclamation by Mayor Michael White declaring March 3, 1995 as Alfonso D'Emilia Day in Cleveland, recognitions in 2006 by Senator George Voinovich, the American Nationalities Movement, and Ohio State Senator Bill Harris, and honored in 2011 as a distinguished member of the Cleveland Community by John Carrol University's Bishop Pilla Program in Italian American Studies.
Alfonso D'Emilia was among the very first people I met in the Italian American Community of Cleveland when I started in the position of Curator for Italian American History at the Western Reserve Historical Society 13 years ago. My colleague, Dr. John Grabowski, had known Alfonso for many years and arranged for us to meet at lunch at Primo Vino in Little Italy along with Gino Colage.
What immediately struck me about Alfonso upon meeting him was his kindness, intelligence and general zest for life. He was so excited to learn that WRHS was beginning a program to preserve the history of Italians in Cleveland, and so very welcoming and encouraging to me in that role, as was his way. He immediately extended an invitation for me to attend the next meeting of Il Cenacolo Italiano di Cleveland at his home in Cleveland Heights.
I attended the meeting of Il Cenacolo at Alfonso's home and met his lovely wife, Mary Jane. I also got my first introduction to the organization and its wonderful members. As I was leaving Alfonso's home, I thanked him for the hospitality. He wished me goodbye with the Italian phrase, "Tante belle cose." I had never heard it before and my comprehension of Italian was not great, so I looked a bit puzzled. Alfonso took my hand and said, "Pamela, dear, it means 'I wish you many beautiful things.'" What a lovely sentiment, I thought. I have used it myself many times since and I always think of Alfonso when I do.
Although Alfonso had retired when I met him, I was lucky enough to hear him speak about several topics at meetings of Il Cenacolo. While my memory of the exact topics he spoke about is fuzzy, I do clearly remember the enthusiasm with which he lectured. The depth and breadth of his knowledge also came through during those talks, which were very thorough and presented without gaps in evidence or reasoning.
Running into Alfonso at various meetings and events, and sometimes at Presti's, was always a treat. Alfonso was always happy to see me and greeted me with a hug and smile. He remembered projects I was working on and my family and would inquire about them with genuine interest. Alfonso was one of the most truly genuine people I have met in my life.
I will always remember Alfonso and the important contributions he made to the city of Cleveland, to the Italian American community and to me personally and professionally. Tante belle cose, Alfonso, as you walk with the angels in heaven.


text:
\section{Introduction}

The nature of the interactions in a turbulent flow has been a long
standing problem. Most investigations are focused on statistically
homogeneous and isotropic turbulence. However, in most cases in nature
this assumption is not necessarily true. Instead, one usually finds
turbulence to be embedded in a large scale flow. In these cases,
turbulence originates from some instability (e.g. convection) and is
not uniformly distributed in space and time. To what extend the flow
is approaching a statistically homogeneous and isotropic state as
smaller and smaller scales are developed is still an open problem.

Due to insufficient computational resources (e.g in astrophysics,
weather and climate prediction, for tokamaks, and in industrial
applications) the computational efforts to study turbulent flows at
these high Reynolds numbers are compelled to resort to modeling of the
smaller scale turbulent fluctuations for which numerical resolution is
not available. However, in order to accurately
model the small scales a good understanding of the impact of the large
scale flow in the small scales is needed, in addition to the knowledge
of what (if any) is the feedback of the small scale turbulent flow
on the large scale structures. This latter process is usually modeled
by an effective viscosity, or other forms of subgrid modeling of the
Reynolds stress tensor. These two types of interactions (from large to small
and from small to large scales, or ``downscaling'' and ``upscaling'')
are often considered to take place only through a local cascade of energy,
although there is evidence that
nonlocal interactions between widely separated scales are also relevant.

Studies of local and nonlocal interactions have been done previously
in direct numerical simulations (DNS), although at moderate spatial
resolutions and Reynolds numbers \cite{Domaradzki88,Domaradzki90,Kerr90,
Yeung91,Okhitani92,Zhou93,Zhou93b,Brasseur94,Yeung95,Zhou96,Kishida99}.
While some of these studies supported the existence of nonlocal
interactions, it was argued by later studies that
their presence was associated with moderate values of the
Reynolds number in the simulations, or that it was linked to the precise
definitions used for transfer functions and the interpretation of
the results. In some cases \cite{Okhitani92,Zhou93,Zhou93b} large
eddy simulations (LES) were also used to extend the range of Reynolds
numbers studied, although if nonlocal interactions are present the impact
of the subgrid model on the transfer is unclear and is a point of study
in itself. Some of these studies
also considered the effect of anisotropic or coherent large scale forcing
(see e.g. Refs. \cite{Yeung91,Yeung95,Zhou96}).
Evidence of nonlocal interactions between large and small scales has
been found also in experiments \cite{Wiltse93,Wiltse98,Carlier01}.
Observations of the persistence of anisotropy at small scales in
experiments at large Reynolds numbers \cite{Sreenivasan97,Shen00}, in
the atmosphere \cite{Stewart69}, and in numerical simulations
(see e.g. Ref. \cite{Biferale01a} and references therein) also suggest
the existence of a direct coupling between disparate scales. The presence
of nonlocal interactions has also been associated in numerical studies
with departures from universality and the development of intermittency
\cite{Laval01}. Recently, a new study at high resolution
\cite{Alexakis05b} presented a detailed analysis of nonlocal interactions
using DNS for large Reynolds numbers, although for a particular and
coherent forcing function.

In this work we examine three different kinds of flows in DNS in
triple periodic boxes, generated by different body forces. We explore
external forcings with and without helicity, and forcings with either infinite
or short correlation times. The energy injection scale is varied, as
well as the viscosity and spatial resolution of the runs, to explore a
wide range of Reynolds numbers and configurations. In all cases, we
observe the presence of nonlocal interactions coexisting with a local
direct cascade of energy. The intensity of the nonlocal interactions
depends on the transfer function studied. For a particular (highly
symmetric) forcing, we also observe a correlation between regions of
large scale shear and strong small scale gradients, supporting previous
studies that linked intermittent effects with interactions with the
large scale flow.

\section{Setup and theory}

\subsection{Equations, code, and simulations}

For an incompressible fluid with constant mass density, the Navier-Stokes
equations are:
\begin{equation}
\partial_t {\bf u} + {\bf u}\cdot \nabla {\bf u} = - \nabla p
    + \nu \nabla^2 {\bf u} +{\bf f} ,
\label{eq:momentum}
\end{equation}
\begin{equation}
\nabla \cdot {\bf u} =0,
\label{eq:incompressible}
\end{equation}
where ${\bf u}$ is the velocity field, $p$ is the pressure divided by
the mass density, and $\nu$ is the kinematic viscosity. Here, ${\bf f}$
is an external force that drives the turbulence. The mode with the largest
wavevector in the Fourier transform of ${\bf f}$ is going to be denoted
as $k_F$ and we are going to refer to $2 \pi k_F^{-1}$ as the forcing
scale. We also define the viscous dissipation wavenumber as
$k_\nu=(\epsilon/\nu^3)^{1/4}$, where $\epsilon$ is the energy injection
rate (as a result, the Kolmogorov scale is $\eta = 2\pi/k_\nu$). A large
separation between the two scales ($k_F^{-1} \gg  k_\nu^{-1}$) is required
for the flow to reach a turbulent state.

In the absence of external forcing and viscosity, the Navier-Stokes
equations in three dimensions have two ideal invariants: the energy
\begin{equation}
E = \frac{1}{2} \int{u^2 d {\bf x}^3} \, ,
\end{equation}
and the helicity
\begin{equation}
H = \frac{1}{2} \int{{\bf u} \cdot \nabla \times {\bf u} \, d{\bf x}^3} \, .
\end{equation}

All the results discussed in the following sections result from analysis
of data from direct numerical simulations of the Navier-Stokes equations.
We solve Eqs. (\ref{eq:momentum}) and (\ref{eq:incompressible}) using a
parallel pseudospectral code in a three dimensional domain of size $2\pi$
with periodic boundary conditions \cite{Gomez05a,Gomez05b}. The pressure
is obtained by taking the divergence of Eq. (\ref{eq:momentum}), using the
incompressibility condition (\ref{eq:incompressible}), and solving the
resulting Poisson equation. The equations are evolved in time using a
second order Runge-Kutta method. The code uses the $2/3$-rule for
dealiasing, and as a result the maximum wavenumber is $k_{max} = N/3$
where $N$ is the number of grid points in each direction. All simulations
presented are well resolved, in the sense that the dissipation wavenumber
$k_\nu$ is smaller than the maximum wavenumber $k_{max}$ at all times.

The Reynolds number is defined as $R_e = UL/\nu$, where $U$ is the
r.m.s. velocity and $L$ is the integral lengthscale of the flow
\begin{equation}
L = 2\pi \frac{\int{E(k) k^{-1} dk}}{\int{E(k) dk}},
\label{eq:integral}
\end{equation}
where $E(k)$ is the energy spectrum. The large scale turnover time
can then be defined as $T=U/L$. We can also introduce the Taylor
based Reynolds number $R_\lambda = U\lambda/\nu$, where the Taylor
lengthscale $\lambda$ is given by
\begin{equation}
\lambda = 2\pi \left(\frac{\int{E(k) dk}}{\int{E(k) k^2 dk}}\right)^{1/2}.
\label{eq:taylor}
\end{equation}

Several simulations were done with different resolutions (from $N=256$ to
$1024$) and kinematic viscosities. Table \ref{table:runs} shows the
parameters for all the runs. The rms velocity in the steady state of all
the runs is $U\approx 1$. To asses the effect of different large scale
stirring forces, we used three expressions for the external force
${\bf f}$. The first expression corresponds to a Taylor-Green (TG) flow
\cite{Taylor37}
{\setlength\arraycolsep{2pt}
\begin{eqnarray}
{\bf f}_{\rm TG} &=& f_0 \left[ \sin(k_F x) \cos(k_F y)
     \cos(k_F z) \hat{x} - \right. {} \nonumber \\
&& {} \left. - \cos(k_F x) \sin(k_F y)
     \cos(k_F z) \hat{y} \right] ,
\label{eq:TG}
\end{eqnarray}}
where $f_0$ is the force amplitude. This expression is not a solution of
the Euler's equations, and as a result small scales are generated fast
when the fluid is stirred with this forcing. The resulting flow has no net
helicity, although regions with strong positive and negative helicity
develop.

In order to study directly the effect of helicity and its transfer
between different scales,
we also did simulations using the Arn'old-Childress-Beltrami (ABC) forcing
{\setlength\arraycolsep{2pt}
\begin{eqnarray}
{\bf f}_{\rm ABC} &=& f_0 \left\{ \left[B \cos(k_F y) +
    C \sin(k_F z) \right] \hat{x} + \right. {} \nonumber \\
&& {} + \left[A \sin(k_F x) + C \cos(k_F z) \right] \hat{y} +
   {} \nonumber \\
&& {} + \left. \left[A \cos(k_F x) + B \sin(k_F y) \right]
   \hat{z} \right\},
\label{eq:ABC}
\end{eqnarray}}
with $A=0.9$, $B=1$, and $C=1.1$. The ABC flow is an eigenfunction of
the curl and an exact solution of the Euler equations. When the flow
is forced using this expression, turbulence develops after an
instability sets in \cite{Podvigina94}. Unlike TG forcing, ABC forcing
injects net helicity into the flow.

The amplitude and phase of these two forcings is kept constant during the
simulations, and as a result the external force has an infinite correlation
time. It is a common practice in studies of isotropic and homogeneous
turbulence to force in Fourier space injecting energy in all modes
in a Fourier shell and changing the phase of each mode with a short
correlation time. To compare with the results of TG and ABC forcing, we
also implemented a random forcing
\begin{equation}
{\bf f}_{\rm RND} = f_0 \sum_{|{\bf k}|=k_F} i {\bf k} \times
    \hat{\bf x} \, e^{i ({\bf k} \cdot {\bf x} + \phi_{\bf k})}
\label{eq:random}
\end{equation}
where the phase $\phi_{\bf k}$ of each mode with wavevector ${\bf k}$ was
changed randomly with a correlation time $\tau_c$ that was taken to be
$\tau_c=0.1 T$.

\begin{table}
\caption{\label{table:runs}Parameters used in the simulations. $N$ is
         the linear grid resolution, ${\bf f}$ the forcing [either Taylor
         Green (TG), Beltrami (ABC) or random (RND)], $k_F$
         the forcing wavenumber, $\nu$ the kinematic viscosity, $R_e$
         the Reynolds number, and $R_\lambda$ the Taylor based Reynolds
         number.}
\begin{ruledtabular}
\begin{tabular}{ccccccc}
Run & $N$  & ${\bf f}$ & $k_F$ &     $\nu$         & $R_e$ & $R_\lambda$ \\
\hline
I   & 256  &  TG       &   2   &$2\times 10^{-3}$  &  675  &     300     \\
II  & 512  &  TG       &   2   &$1.5\times 10^{-3}$&  875  &     350     \\
III & 1024 &  TG       &   2   &$3\times 10^{-4}$  & 3950  &     800     \\
IV  & 256  &  ABC      &  10   &$2.5\times 10^{-3}$&  275  &     230     \\
V   & 256  &  ABC      &   3   &$2\times 10^{-3}$  &  820  &     360     \\
VI  & 512  &  ABC      &   3   &$6.2\times 10^{-4}$& 2520  &     670     \\
VII & 1024 &  ABC      &   3   &$2.5\times 10^{-4}$& 6200  &    1100     \\
VIII& 256  &  RND      &   1   &$1.5\times 10^{-3}$& 2030  &     650     \\
\end{tabular}
\end{ruledtabular}
\end{table}

\subsection{Scale interactions and transfer}
To investigate the interactions between different scales we split the
velocity field into spherical shells in Fourier space of unit width,
{\it i.e.}\  $\mathbf{v}=\sum_K {\bf v}_K$ where ${\bf v}_K$ is a filtered velocity
field such that only the Fourier modes in the shell $K\le|k|<K+1$
(from now on called shell $K$) are kept.
From equation
(\ref{eq:momentum}), the rate of energy transfer $T_3(K,P,Q)$ (a
third-order correlator) from energy in shell $Q$ to energy in shell $K$
due to the interaction with the velocity field in shell $P$ is defined
as usual \cite{Kraichnan71,Lesieur,Alexakis05,Alexakis05b} as:
\begin{equation}
T_3(K,P,Q) = -\int \mathbf{v}_K \cdot (\mathbf{v}_P \cdot \nabla) \mathbf{v}_Q d{\bf x}^3 \ .
\label{trans_eq3}
\end{equation}
Note that this term does not give information about the energy the
shell $P$ receives or gives to the shells $K$ and $Q$.
The computation of $T_3$ for the three shells $K$, $P$, and $Q$
from 1 up to a wavenumber $K_{max}$ requires $\sim K_{max}^3 N^3$
operations and is therefore demanding in computer resources. For example,
in the $1024^3$ simulations, to compute $T_3(K,P,Q)$ for all wavenumbers
up to $K_{max}=80$, it takes as much computing time as
evolving the hydrodynamic code for two turnover times.

If we sum over the middle wave number $P$ we obtain the total energy
transfer $T_2(K,Q)$ from shell $Q$ to shell $K$:
\begin{equation}
T_2(K,Q) = \sum_P T_3(K,P,Q) = -\int \mathbf{v}_K \cdot (\mathbf{v} \cdot \nabla)
\mathbf{v}_Q d{\bf x}^3 \ .
\label{trans_eq2}
\end{equation}
Positive transfer implies that energy is transfered from shell $Q$ to $K$,
and negative from $K$ to $Q$; thus, both $T_3$ and $T_2$ are antisymmetric
in their $(K,Q)$ arguments (see Ref. \cite{Alexakis05}). $T_2(K,Q)$ gives
information on the shell-to-shell energy transfer between $K$ and $Q$,
but not about the amplitude of the triadic interactions
themselves.

If we further sum over the wave number $Q$ we obtain the transfer
\begin{equation}
T_1(K) = \sum_Q T_2(K,Q) = -\int \mathbf{v}_K \cdot (\mathbf{v} \cdot \nabla) \mathbf{v} d{\bf x}^3
\label{trans_eq1}
\end{equation}
that expresses the rate the shell $K$ receives energy from the velocity
field (all shells).

The energy flux is reobtained from these transfer functions as
\begin{equation}
\Pi(k) = -\sum_{K=0}^k T_1(K)=-\int \mathbf{v}_K^< (\mathbf{v}\cdot \nabla )\mathbf{v}_K d{\bf x}^3
\label{eq:flux}
\end{equation}
where the notation
\begin{equation}
\mathbf{v}_K^<=\sum_{K'=0}^K \mathbf{v}_{K'} \,\,\, \mathrm{and} \,\,\,
\mathbf{v}_K^>=\sum_{K'=K+1}^\infty \mathbf{v}_{K'}
\end{equation}
is used (see Ref. \cite{Frisch}).

We can further define the flux of energy at some given scale due to
the interactions with some other scale as
\begin{equation}
\Pi_P(k) = -\sum_{K=0}^k \sum_Q T_3(K,P,Q)=
-\int \mathbf{v}_K^< \cdot (\mathbf{v}_P \cdot \nabla) \mathbf{v}^> d{\bf x}^3 \,\, ,
\end{equation}
which expresses the flux of energy at the scale $K^{-1}$ due to the
interactions with the scale $P^{-1}$
or in other words the energy flux at the scale $K^{-1}$ if the velocity
field was advected just by the velocity field at scale $P^{-1}$.
As will be discussed later, the
question of locality depends on which of the different transfer
functions or flux is under investigation.

It is worth noticing that through the study of the three functions
$T_1$, $T_2$, $T_3$, and the fluxes $\Pi_P$, we can measure the nature
and intensity of the interactions directly,
and therefore we avoid
introducing a non-locality parameter as was done e.g. in Refs.
\cite{Zhou93,Zhou93b}.

The total energy balance equation for a given shell is written as
\begin{equation}
\partial_t E(K) = T_1(K)+ \nu {\mathcal D(K)} + {\mathcal F}(K)
\end{equation}
where we have also introduced the dissipation function
\begin{equation}
\nu {\mathcal D}(K) \equiv  \nu   \int |{\bf \nabla u}_K|^2  d{\bf x}^3 ,
\end{equation}
and the energy injection rate to the velocity field through the forcing term
\begin{equation}
{\mathcal F}(K) \equiv \int { \bf f} \cdot {\bf u}_K  \, d{\bf x}^3 \,.
\end{equation}

Finally, we will also investigate the transfer of helicity
among different scales. Taking the inner product of the
Navier-Stokes equation (\ref{eq:momentum}) with the vorticity at a scale $K^{-1}$
($\mathbf{w}_K=\nabla \times \bu_K$) and adding the inner product of the velocity at
the same scale $K^{-1}$ with the $curl$ of (\ref{eq:momentum})
and space averaging we obtain:
\begin{eqnarray}
\partial_tH(K)=&\sum_Q \int \mathbf{w}_K \cdot (\bu \times \mathbf{w}_Q) d{\bf x}^3+
    \nonumber \\
&\int \mathbf{w}_K \cdot F d{\bf x}^3 +\nu \int \mathbf{w}_K \cdot \nabla \times
    \mathbf{w}_K d{\bf x}^3.
\end{eqnarray}
Each term of the sum in the first term in the $r.h.s$
of the equation above can be written as:
\begin{equation}
T_H(K,Q) = \int \mathbf{w}_K \cdot (\bu \times \mathbf{w}_Q) d{\bf x}^3 \ .
\label{trans_eqh}
\end{equation}
Note that the anti-symmetric property $T_H(K,Q)=-T_H(Q,K)$ holds for
$T_H(K,Q)$ {\it i.e.}\  the rate the shell $K$ is gaining helicity from the
interaction with the field $\mathbf{w}_Q$ and $\bu$ is equal to the rate the
shell $Q$ is loosing helicity through the same interaction. This allows
us to interpret $T_H(K,Q)$ as the transfer of helicity from the
scale $Q^{-1}$ to the scale $K^{-1}$ (see \cite{Alexakis06} for the
equivalent definition of the magnetic helicity transfer).

\section{\label{sec:runIII}Influence of the large scale flow on turbulence}

We discuss in this section properties of Run III. First we describe the
geometry of the resulting flow and we present general statistical results,
such as probability density functions and structure functions. Then, we
analyze in detail the scale interactions using the formulation discussed
in the previous section. Preliminary results for this flow were presented
in \cite{Alexakis05b}. The results discussed here will be used as a
reference to compare with the rest of the simulations in the following
sections.

\subsection{Statistical properties, structure functions, and scaling exponents}

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig1}}
\caption{(Color online) Rendering of enstrophy density in a
         $1024 \times 64 \times 1024$
         slice of Run III. Bands of strong (1,2,3,4) and weak (A,B,C,D)
         shear in the external Taylor-Green force are indicated. The bars
         at the bottom show respectively the integral, Taylor, and
         dissipation scales.}
\label{fig:render}
\end{figure}

Run III is a $1024^3$ simulation using TG forcing. After reaching a
turbulent steady state, the simulation was continued for 10 turnover
times. The expression of the forcing has several spatial symmetries, and
some of these symmetries are recovered in the flow in a statistical
sense. Figure \ref{fig:render} shows a rendering of enstrophy density in a
thin slice of $1024 \times 64 \times 1024$ grid points. Thin and elongated
structures (vortex tubes) can be identified. As observed in previous
studies, these structures are characterized by one large lengthscale
in one direction (ranging from the Taylor to the integral scale) and
a short lengthscale (close to the viscous dissipation scale) in the other
two directions.

\begin{figure}
\centerline{\includegraphics[width=7cm]{fig2}}
\caption{(Color online) Zoom on a region with large enstrophy density
         showing field lines inside
         (above) and in its surrounding (below) in Run III. Velocity field
         lines are helical inside and in the vicinity of the tube. Note the
         merging of two tubes in the south west corner. The side of the region
         shown is approximately one tenth of the side of the total
         domain.}
\label{fig:tube}
\end{figure}

Since $k_F=2$, there are four planes in the $z$ direction where the
external force is zero (see Eq. \ref{eq:TG}) and the shear is maximum.
In the steady state turbulent flow, these planes can be easily recognized
since four bands where the vorticity is stronger are formed around them.
Regions with less (or weaker) vortex tubes separate these planes,
regions centered around the planes where the large scale shear has
a minimum. Together, these two sets of regions
form a large scale pattern that is observed to persist for long times,
giving four ``quiet'' stripes and four stripes of ``stronger" turbulence
(the boundary between these regions is not well defined and fluctuates
in space and time, see Fig. \ref{fig:render}). Considering the persistence
of these statistical
symmetries of the flow, the velocity increments for this Run will be
computed for displacements {\it only} in the $x$-$y$ plane.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig3}}
\caption{Energy spectrum compensated by $k^{-5/3}$ for Run III. The
         inset shows the energy flux.}
\label{fig:spectrum_runIII}
\end{figure}

When individual vortex tubes are studied, it is seen that the
flow inside and surrounding the vortex tube is helical, as noted before by several authors
\cite{Tsinober83,Moffatt85,Moffatt86,Levich87,Farge01}. Figure
\ref{fig:tube} shows a rendering of a small region in the domain
(approximately one tenth of the box). A region of large enstrophy
density is indicated by dark (blue) colors, and field lines inside
and in the vicinity of the enstrophy containing region are shown.
In this region and its surroundings, field lines are helical, while
the flow far from the region is not. This feature has been verified
for a large set of vortex tubes in the complete domain. It has been
claimed in the past \cite{Moffatt92,Tsinober} that the development of these
helical structures in a turbulent flow can lead to the depletion of
nonlinearity and a quenching of local interactions. We will come back
to this issue in the following sections.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig4}}
\caption{Probability density functions of transverse velocity increments
         for Run III, for increments $l=20 \eta$ ($\cdots$), $l=10 \eta$
         ($- - -$), $l=4 \eta$ ($-\cdot-$), $l=2 \eta$ ($-\cdots-$), and
         $l=\eta$ (solid line), where $\eta$ is the Kolmogorov dissipation
         scale.}
\label{fig:pdf_runIII}
\end{figure}

In spite of the presence of the large scale pattern, many features often
associated with isotropic and homogeneous turbulence can be observed
in this simulation. Figure \ref{fig:spectrum_runIII} shows the
angle-averaged Fourier energy spectrum and energy flux. An inertial
range with constant energy flux is observed, together with a
Kolmogorov-like scaling and a bottleneck as the dissipative range is
reached. When probability density functions (pdfs) of transverse
velocity increments
\begin{equation}
\delta v_\perp ({\bf x},l) = \hat{\bf r} \times \left[ {\bf v}({\bf x} +
    l\hat{\bf r}) - {\bf v}({\bf x}) \right],
\label{eq:incrementperp}
\end{equation}
are computed in the whole domain (Fig. \ref{fig:pdf_runIII}), we observe
distributions close to Gaussian for large increments, and the development
of exponential and stretched exponential tails as the increment $l$ is
decreased.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig5}}
\caption{Probability density functions of transverse velocity increments
         with $l=10 \eta$ for Run III. The solid line corresponds to the
         whole domain, dashed lines to regions 1 to 4 (strong shear),
         and dotted lines to regions A to D (weak shear; see Fig.
         \ref{fig:render}). Notice the weak but systematic differences:
         the four dashed lines (some of them overlapping) have slower decaying
         tails and the dotted lines have faster decaying tails.}
\label{fig:pdf10eta}
\end{figure}

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig6}}
\caption{Probability density functions of transverse velocity increments
         with $l=2 \eta$ for Run III. Same labels as in Fig.
         \ref{fig:pdf10eta}. Note that the differences between quiet
         and strong regions appear even more systematic at this scale than in Fig.
         \ref{fig:pdf10eta}.}
\label{fig:pdf2eta}
\end{figure}

Figures \ref{fig:pdf10eta} and \ref{fig:pdf2eta} show pdfs of transverse
velocity increments with $l=10 \eta$ and $2 \eta$ respectively, but
discriminating between bands of strong and weak shear (as indicated in Fig.
\ref{fig:render}). Bands with strong shear (1,2,3, and 4) have
slightly but systematically stronger
tails (i.e. a larger probability of strong gradients) than regions A,B,C,
and D. Note this behavior is true for each individual region, and the
difference increases as the increment $l$ is decreased. This confirms
statistically what can be inferred from Fig. \ref{fig:render}: regions
of strong shear have a larger density of vortex tubes and stronger
gradients.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig7}}
\caption{First order structure function $S_1(l)$ for Run III. The solid
         line corresponds to the whole domain, dashed lines to regions
         1,2,3, and 4, and dotted lines to regions A,B,C, and D. The
         Kolmogorov scaling and the linear slope corresponding
         to a smooth flow are shown as a reference. The arrow indicates
         the Kolmogorov scale $\eta$. Notice again the systematic
         differences between the two regimes identified in Fig. 1.}
\label{fig:S1}
\end{figure}

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig8}}
\caption{Second order structure function $S_2(l)$ for Run III. Labels
         are as in Fig. \ref{fig:S1}.}
\label{fig:S2}
\end{figure}

We can also compute longitudinal velocity increments
\begin{equation}
\delta v_\parallel ({\bf x},l) = \hat{\bf r} \cdot \left[ {\bf v}({\bf x} +
    l \hat{\bf r}) - {\bf v}({\bf x}) \right] .
\label{eq:incrementpar}
\end{equation}
Figures \ref{fig:S1} and \ref{fig:S2} show respectively the second and
third order longitudinal structure functions, where the structure function
of order $p$ is defined as
\begin{equation}
S_p (l) = \left< \delta v_\parallel^p ({\bf x},l) \right> .
\label{eq:structure}
\end{equation}
At scales smaller than the dissipation scale the field is smooth and
the structure function of order $p$ scales as $l^p$. Kolmogorov's 1941
theory of turbulence predicts in the inertial range a scaling
$S_p (l) \sim l^{p/3}$, although corrections due to intermittency are
known and in general $S_p (l) \sim l^{\zeta_p}$, where $\zeta_p\not= p/3$ are
the scaling exponents.

As with pdfs of velocity increments, a clear trend separating regions of
strong and weak shear is observed. In the range of scales corresponding to
the inertial range, the four regions with strong shear (regions 1,2,3 and 4)
show a larger slope than the four regions with weak shear. The slope of the
structure function computed in the whole box lies between these two
values. Considering the results from both pdfs and structure functions,
a correlation is observed between small scale gradients and large scale
shear. Note that a correlation between stronger tails in the pdfs
of velocity increments and vortex tubes has already been observed in
\cite{Levi01,Farge01}. Here, a correlation between these quantities with
large scale shear is further observed, in agreement with Ref. \cite{Laval01}
that suggests intermittency is related with interactions with the large
scale flow.

\subsection{Interactions, energy transfer, and flux}

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig9}}
\caption{(Color online) Above: amplitude of triadic interactions
         as given by $T_3(K,P,Q)$
         at $Q=40$ for Run III. Solid lines correspond to positive contour
         levels, and dotted lines to negative values. The dark (red) shaded
         region indicates shells in $P$ close to the forcing shell, while
         the light (orange) shaded region indicates shells in $P$ outside
         the local octave. Below: total shell-to-shell energy transfer
         $T_2(K,Q)$, transfer $T_2^{\textrm LS}(K,Q)$ due to interactions
         with $P$ close to the forcing shell (dashed line), and difference
         between these two transfers (dash-dotted line).}
\label{fig:triads}
\end{figure}

Figure \ref{fig:triads} shows the functions $T_3(K,P,Q)$ and $T_2(K,Q)$
evaluated at $Q=40$ in the turbulent steady state of Run III. When triadic
interactions between shells are studied, the strongest interactions are
with the shell $P=3$, where the large scale forcing is. Local
interactions between Fourier shells ($K \sim P \sim Q$) are two orders
of magnitude smaller than nonlocal interactions with $P \sim k_F$.
As a result, when $T_3(K,P,Q)$ is summed over all $P$ to obtain the
shell-to-shell energy transfer $T_2(K,Q)$, the energy is observed to
be transfered to small scales locally between shells, but with a step
proportional to $k_F$. In $T_2(K,Q)$, the negative peak at
$K \sim Q-k_F$ means energy is transfered from this shell to the shell
$Q$, while the positive peak at $K \sim Q+k_F$ indicates energy is
transfered to this shell from the shell $Q$.

The function $T_3(K,P,Q)$ was studied in the steady state of this run
also for $Q=10$ and $Q=20$, obtaining the same quantitative results: a
dominance (when compared in amplitude) of triadic interactions with the
large scale flow over local triadic interactions.
The $T_2(K,Q)$ function for all values of $Q$ between $10$ and $80$ also
peaks at $K\approx Q \pm k_F$ (see Ref. \cite{Alexakis05b}). These peaks
at fixed values of $K-Q$ in $T_2(K,Q)$ are just the signature of the
strong triadic interactions with the large scale forcing. If we compute
the shell-to-shell transfer between shells $K$ and $Q$ due only to
interactions with the large scale flow
\begin{equation}
T_2^{\textrm LS} (K,Q) = \sum_{P=0}^6 T_3(K,P,Q) \, ,
\label{eq:T2LS}
\end{equation}
we obtain most of the shell-to-shell transfer (see Fig. \ref{fig:triads}).
The remaining transfer $T_2(K,Q)-T_2^{\textrm LS}(K,Q)$ still peaks at
larger wavenumbers, although it still does not peak at $K \sim Q = 40$.
This is due to nonlocal interactions outside the octave
band not related with the large scale forcing (light shaded region in
Fig. \ref{fig:triads}). Note that the definition of the range summed over
$P$ in Eq. (\ref{eq:T2LS}), or of octave bands (here defined as
$Q/2 \lesssim K \lesssim 2Q$) is somewhat arbitrary.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig10}}
\caption{Total energy flux $\Pi(k)$ in Run III (solid line), flux
         $\Pi^{\textrm LS}(k)$ due to interactions with the large scale
         flow (dashed line), and nonlocal flux $\Pi^{\textrm NL}(k)$.}
\label{fig:fluxes_runIII}
\end{figure}

As previously mentioned, while individual triadic interactions as described
by $T_3(K,P,Q)$ are dominantly nonlocal, the shell-to-shell transfer
$T_2(K,P,Q)$ describes a local transfer of energy although through
interactions with the large scale flow. Clearly when we sum over shells,
the larger number of modes in the small scales start dominating over the
large scale modes. Summing further over $K$ and $Q$, from Eq.
(\ref{eq:flux}) we can obtain the total energy flux. In analogy with Eq.
(\ref{eq:T2LS}) we can also define the flux due to interactions with the
large scale flow (dark shaded region in Fig. \ref{fig:triads})
\begin{equation}
\Pi^{\textrm LS}(k) = \sum_{P=0}^{6} \Pi_P(k) \, ,
\end{equation}
and the nonlocal flux due to interactions outside an octave band (light
shaded regions in Fig. \ref{fig:triads})
\begin{equation}
\Pi^{\textrm NL}(k) = \sum_{P=7}^{k/2} \Pi_P(k) \, ,
\end{equation}
where $\Pi_P(k)$ is defined in Eq. (15).

Figure \ref{fig:fluxes_runIII} shows these three fluxes in the steady
state of Run III. The large scale flow is only responsible for a small
(but not insignificant) fraction of the total energy flux
($\Pi^{\textrm LS} \simeq 0.2 \Pi$).
Furthermore although the amplitude of the nonlocal flux
$\Pi^{\textrm NL}(k)$ depends on the definition of an octave band, it is
remarkable that it peaks at wavenumbers close to the peak of the
bottleneck in the energy spectrum (see Fig. \ref{fig:spectrum_runIII}).
This gives a direct confirmation that the bottleneck is due to the
depletion of the energy transfer due to local interactions with
$K \sim P \sim Q$. These interactions are inhibited because of the
presence of a numerical and/or viscous cutoff in wavenumber, and
nonlocal triads become dominant
\cite{Herring82,Falkovich94,Lohse95,Martinez97}. Indeed, in Ref.
\cite{Herring82} it was shown using the eddy-damped quasinormal Markovian
(EDQNM) approximation that the bottleneck disappears if nonlocal
interactions are excluded from the computation. Note that the spurious
drop of the fluxes $\Pi^{\textrm LS}$ and $\Pi^{\textrm NL}$ at $k=80$
in Fig. \ref{fig:fluxes_runIII} is because the functions $T_3(Q,P,Q)$
and $T_2(K,Q)$ were only computed up to this wavenumber.

Therefore there seems to be a hierarchy concerning the importance of
nonlocality when one investigates the transfer functions.
At the most basic level when one investigates the transfer function $T_3$
the interactions with the large scale flow are the strongest ones.
When averaged over the middle wave number the resulting energy transfer $T_2$
becomes local but not self-similar with a maximum at $K-Q\sim k_F$. When
averaged further to obtain the energy flux, the non local interactions with
the large scale flow are only responsible for  20\% of the flux.

\section{\label{sec:Reynolds}Scaling with Reynolds}

In this section we discuss results from runs I, II, and III.
The three runs are forced using Eq. (\ref{eq:TG}), and the only parameters
changed between the runs are the spatial resolution and the kinematic
viscosity. This investigation allows us to study how the results about
the non-local interactions change as we increase the Reynolds number.
Qualitatively the functions $T_3(K,P,Q)$ and $T_2(K,Q)$ are similar for
all three runs: the strongest interactions [$T_3(K,P,Q)$] are with
the large scale flow and the energy transfer [$T_2(K,Q)$] is local with
the two peaks at $K-Q \sim k_F$.
For this reason we focus here on the study of the local
and nonlocal flux of energy as the Reynolds number is changed.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig11}}
\caption{Scale variation of the flux ratio $\Pi^{\textrm LS}/\Pi$ for
         increasing Reynolds numbers; Runs I (dash-dot line),
         II (dash line), and III (solid line). Note the presence of a wavenumber range
         where $\Pi^{\textrm LS}/\Pi$ becomes constant only for the
         highest resolution run.}
\label{fig:fluxratio}
\end{figure}

In Figure \ref{fig:fluxratio} we show the ratio of the flux
$\Pi^{\textrm LS}(k)$ to the total flux $\Pi(k)$. As discussed in the
previous section, the flux due to interactions with the large flow gets
diminished as we sum over more and more modes. As a result, in the
inertial range of Run III interactions with the large scale flow are
only responsible for $\sim 20\%$ of the energy flux. In the smaller
Reynolds number runs (I and II), the amount of flux due to interactions
with the large scale flow is larger. This implies that as we increase
the Reynolds number the fraction of the energy flux in the inertial
range due to interactions with the large scale flow decreases.
However, note that in Run III a region in the inertial range where
the ratio $\Pi^{\textrm LS}/\Pi$ is approximately constant is observed.
This region with constant ratio $\Pi^{\textrm LS}/\Pi$ is not present
in the simulations with lower Reynolds number.
Therefore DNS with even higher Reynolds number than what is accomplished
here are needed
to determine how does this ratio scale with the Reynolds number.

\section{Effects of different forcing functions}

\subsection{Forcing expression and correlation time}

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig12}}
\caption{Compensated energy spectrum $k^{5/3} E(k)$ (solid line) and
         helicity spectrum $k^{5/3} H(k)/k_F$ (dashed line) in the
         steady state of Run VII, where $k_F$ is the forcing wavenumber.}
\label{fig:spectrum_runVII}
\end{figure}

Nonlocal interactions discussed in the previous section were observed
for runs with large scale non-helical forcing with infinite correlation
time. It is of interest to know how much of these results translate to
other kind of forcings. In this section we compare results for runs
III, VII and VIII. Run VII is a $1024^3$ simulation with constant helical
forcing, and as runs I-III, it also displays a well defined large scale
flow (for a description of the ABC flow, see e.g. Ref. \cite{Childress}).
In Run VIII the phases of the forcing function are changed randomly with
a short correlation time. Although the resolution in this run is smaller
($256^3$), we will focus here on the effect of the forcing, since a
systematic study of changing the Reynolds number for fixed forcing
was presented in the previous section.

Figure \ref{fig:spectrum_runVII} shows the energy and helicity
spectra compensated by a Kolmogorov law in the turbulent steady state of Run VII.
As proposed in Ref. \cite{Brissaud73} and
observed in several simulations
\cite{Andre77,Borue97,Chen03,Chen03b,Gomez04} the spectrum of helicity
follows a Kolmogorov law and is proportional to $k_F E(k)$. The transfer
of helicity will be discussed in Section \ref{sec:helicity}.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig13}}
\caption{Scaling exponents of velocity structure functions $\zeta_p$
         as a function of order $p$
         in Runs III (TG forcing) and VII (ABC forcing). The Kolmogorov
         K41 scaling \cite{Kolmogorov41}  is indicated, as well as the She-Leveque (SL)
         prediction \cite{She94}. Note the difference between TG and ABC.}
\label{fig:scaling}
\end{figure}

Figure \ref{fig:scaling}
shows the scaling exponents $\zeta_p$ of the longitudinal structure
functions, computed for runs III and VII. The extended self-similarity
(ESS) hypothesis \cite{Benzi93,Benzi93b} was used to compute the anomalous
exponents, which show similar behavior for both runs. However,
the resulting exponents for $p>3$ from Run VII (ABC forcing) are slightly
smaller than the exponents from Run III (TG forcing); for example,
 in Run III $\zeta_4 = 1.2737 \pm 0.0005$ and
$\zeta_8 = 2.136 \pm 0.006$, while in Run VII $\zeta_4 = 1.2647 \pm 0.0005$
and $\zeta_8 = 2.026 \pm 0.008$.
Note that the difference
between the scaling exponents for TG and ABC at order $p=4$, though small, is
nevertheless more than order of magnitude larger than the error arising from
measuring them by virtue of the ESS hypothesis which leads to negligible
errors.

It should be noted that it is unclear whether
this observed difference is due to
problems linked with using the ESS methodology itself, or if it indicates
a departure from universality between
these two flows because of loss of homogeneity, as exemplified in Fig. 1.
Since the two forcing functions studied here are both to some degree
anisotropic and inhomogeneous, subleading contributions due to departures
from full symmetry could be responsible for
this discrepancy. In this context, a decomposition of the structure
functions into their isotropic and
anisotropic components \cite{Biferale01a,Biferale01b,Biferale02} could
help to asses the degree of universality of each component.
These points will necessitate further study and better resolved flows.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig14}}
\caption{Shell-to-shell energy transfer function $T_2(K,Q)$ for Run VII.
         The superimposed curves correspond to different values of $Q$,
         with $Q\in [10,80]$.}
\label{fig:T2_runVII}
\end{figure}

The shell-to-shell energy transfer $T_2(K,Q)$ for Run VII is shown in Fig.
\ref{fig:T2_runVII}. As for non-helical forcing (see Sec. \ref{sec:runIII}
and Ref. \cite{Alexakis05})
the shell-to-shell transfer function peaks at $K \approx Q \pm k_F$,
indicating the transfer of energy is local but mediated by nonlocal
interactions with the large scales. If we sum over $K$ and $Q$ to
obtain the energy flux, we reobtain the results discussed in Sec.
\ref{sec:Reynolds} for Run III: most of the energy flux is local,
although in the inertial range $\sim 20\%$ of the total flux is due
to interactions with the large scale forcing.

The transfer function $T_2(K,Q)$ was also computed in the turbulent
steady state of Run VIII (see Fig. \ref{fig:T2_scale}.a). In this run
the phases of the external force are changed randomly with a short
correlation time. It is noteworthy that even in this case with
isotropic and random forcing, evidence is found of nonlocal
interactions with the large forcing scale: the $T_2(K,Q)$
function peaks at $K \approx Q \pm k_F$ for all wavenumbers $Q$ studied
(note that in this run $k_F=1$). Fig. \ref{fig:T2_scale} is discussed further below.

\subsection{Forcing scale}

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig15}}
\caption{Shell-to-shell energy transfer function $T_2(K,Q)$ for runs with
         different forcing wavenumbers $k_F$ as indicated by the vertical
         lines; Run VIII with $k_F=1$ (a), Run V with $k_F=3$ (b), and
         Run IV with $k_F=10$ (c). In each panel, the several curves
         correspond to different values of $Q$ in the inertial and
         dissipative ranges. Note that for each run, the peak of energy
         transfer is centered close to $k_F$. }
\label{fig:T2_scale}
\end{figure}

Through all this work we have shown transfer functions indicating that the
energy is locally transfered between scales but with a fixed step that
we associated with the forcing scale. In this subsection we compare
the transfer function $T_2(K,Q)$ in simulations forced at different
wavenumbers.

Figure \ref{fig:T2_scale} shows $T_2(K,Q)$ for several
values of $Q$ computed in the steady state of runs VIII ($k_F=1$),
IV ($k_F=2$), and V ($k_F=10$). A clear correlation is observed
between the position of the peaks ($|K-Q|$) in the shell-to-shell
transfer function, and the forcing wavenumber $k_F$. In the last
case, the peaks at $|K-Q|$ slightly smaller than $k_F$ could be
associated with low Reynolds number effects.

\section{\label{sec:helicity}Helicity transfer}

The transfer of helicity is readily studied in Run VII, since the
external forcing injects maximum helicity and all scales are dominated
by the same sign of helicity. We will make no attempt here to separate
the different signs of helicity in the simulations (see e.g. Refs.
\cite{Waleffe91,Chen03,Chen03b}).

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig16}}
\caption{The spectrum of relative helicity $H(k)/kE(k)$
         compensated by $k^{-1}$ for run VII. Note that the region where the
         spectrum is flat indicates a power law of $k^{-1}$ in the relative
         helicity. At large wavenumber, a $k^{1/2}$ slope is shown only
         as a reference. The increase at large $k$ is indicative of an
         excess of relative helicity in the small scales.}
\label{fig:relative}
\end{figure}

Before discussing the transfer of helicity, it is of interest to study
spectral properties of helical flows. As previously discussed, the spectrum
of helicity follows an approximate $k^{-5/3}$ law (see Fig.
\ref{fig:spectrum_runVII}). As a result, the relative helicity $H(k)/E(k)k$
follows in the inertial range a $k^{-1}$ slope
\cite{Brissaud73,Andre77,Lesieur} (i.e. small scales are less helical
than large scales).
It was predicted in Refs. \cite{Ditlevsen01,Ditlevsen01b}
that the dissipation scale of helicity should be larger than the energy
dissipation scale, giving as a result a drop in the spectrum of relative
helicity faster than $k^{-1}$ for small scales. This argument would be
in agreement with the idea that small scales slowly recover the mirror
symmetry broken by the injection of helicity in the large scales. However,
previous simulations \cite{Gomez04} and this high resolution run both suggest that
there is an excess of relative helicity in the small scales, when compared
with the $k^{-1}$ drop. Note that this slower than predicted recovery of
symmetries in the small scales is also in agreement with the slower than
expected recovery of isotropy observed in experiments
\cite{Sreenivasan97,Shen00}.

Figure \ref{fig:relative} shows $H(k)/E(k)$, i.e. the spectrum of relative
helicity $R(k)=H(k)/kE(k)$ compensated by $k^{-1}$.
A scaling of $k^{-1}$ for the relative helicity $R(k)$ thus corresponds
to a flat spectrum in Fig. 16,
as observed through the inertial range up to $k\sim 20$. However, at small
scales the compensated spectrum of relative helicity grows,
possibly as $k^{1/2}$ or steeper, indicating that
the spectrum of helicity $H(k)$ at small scales is dropping slower than
the spectrum of energy $E(k)$ (see also Fig. \ref{fig:spectrum_runVII}).
This can be associated with the presence of vortex tubes, that are
known to be helical (see Fig. \ref{fig:tube}).

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig17}}
\caption{Energy spectrum compensated by $k^{-5/3}$ for Runs V (dashed
         line), VI (dotted line), and VII (solid line). All runs are for
         ABC forcing at different Reynolds numbers. Note the wavenumber
         on the x-axis has been divided by the dissipation wavenumber to
         make all dissipation ranges coincide. A slope of $1/3$ close to
         the bottleneck (corresponding to a power law of $k^{-4/3}$ in
         the energy spectrum) is shown as a reference.}
\label{fig:bottleneck}
\end{figure}

The effect of helicity, and indirectly the effect of nonlocal interactions
which can be modeled for example
through the introduction of a second time-scale in the problem in order to
distinguish between the eddy turn-over time and a helical characteristic
time, has also been used to explain the development of the bottleneck
effect that occurs at the onset of the dissipative range, see e.g.
\cite{Kurien04}. Ref. \cite{Kurien04} predicts a $k^{-4/3}$ energy
spectrum for the bottleneck, that we found compatible with our spectra
for runs V-VII (Fig. \ref{fig:bottleneck}). We also observe a $k^{-4/3}$
range close to the bottleneck in the Taylor-Green flow that has no net
helicity. Note that the argument in Ref. \cite{Kurien04} is based
on the presence of non zero helicity locally. The transition from the
inertial range to the bottleneck seems to be dependent on the Reynolds
number. It is worth noting that while at resolutions of $256^3$ mostly
a bottleneck is observed, in the $1024^3$ runs a short Kolmogorov-like
scaling is found before the bottleneck takes place. The origin of the
bottleneck based on the dominance of nonlocal interactions in the
dissipative range has also been investigated
\cite{Herring82,Falkovich94,Lohse95,Martinez97} as we discussed in
Sec. \ref{sec:runIII}, and is independent of the presence of helicity.
These arguments are not mutually exclusive, as local in space generation
of helicity in the small scales can also quench local interactions between
eddies of comparable sizes.
However, the prediction in Ref. \cite{Falkovich94} for the spectral
shape of the bottleneck is in disagreement with the spectra obtained
in all simulations here. These points deserve further study at higher
resolution if one is to be able to distinguish between the different
ranges that may be occurring.

\begin{figure}
\centerline{\includegraphics[width=8.3cm]{fig18}}
\caption{Shell-to-shell helicity transfer $T_H(K,Q)$ in Run VII, for
         all values of $Q$ from $10$ to $80$.}
\label{fig:T2H}
\end{figure}

Finally, we discuss the shell-to-shell helicity transfer $T_H(K,Q)$
in Run VII (Fig. \ref{fig:T2H}). As for the transfer of energy,
the helical transfer function peaks at $K \approx Q \pm k_F$, although this
transfer is noisier. The helicity cascades directly to smaller scales
as the energy, confirming previous studies
\cite{Borue97,Chen03,Chen03b,Gomez04}.
The helicity is not a positive definite quantity, and as a result both
signs of the transfer can indicate a direct cascade depending on the
sign of the helicity itself. Moreover, as we study smaller scales and the
relative helicity decreases, both signs have more predominance
increasing the noise observed in $T_H(K,Q)$.

\section{Conclusions}

We start by summarizing our results. The energy transfer, triadic
interactions, and statistical flow properties were studied in
several high resolution DNS with periodic boundaries. The spatial
resolution ranged from $256^3$ to $1024^3$. The Reynolds number based
on the integral scale spanned values from $R_e \approx 275 $ to $6200$,
while the Taylor-based Reynolds number varied between
$R_\lambda \approx 230$ and $1100$. Three forcing functions were used,
two coherent functions with zero and maximum net helicity respectively,
and a random forcing with a short correlation time. The energy injection
scale was also varied to study its impact on the energy transfer.

Most statistical studies were done for Run III, a high resolution
simulation with $R_\lambda \approx 800$ and TG forcing. The forcing and
the resulting flow have spatial symmetries that allow us to identify
planes with strong and weak large scale shear easily. While in the
whole domain the standard results were reobtained (e.g. exponential
and stretched exponential tails in the pdfs of velocity increments,
and anomalous scaling of the structure functions), when studying
individual regions a correlation between large scale shear and small
scale gradients was found. In regions of strong shear, the tails of
the pdfs of velocity increments are stronger than in regions of weak
shear, even at scales as small as twice the dissipation scale. Also,
structure functions show a slightly larger slope in regions of strong
large scale shear, i.e. a larger departure from a Kolmogorov $p/3$ self-similar scaling.

A correlation between stronger tails in the pdfs of velocity
increments and the presence of vortex tubes has already been observed
in different decompositions of the flow \cite{Levi01,Farge01}. Here,
we observed a correlation between these quantities and the large scale
shear. The correlation was observed to be persistent even at scales as
small as the dissipation scale. Small differences were also observed
in the anomalous scaling of the structure functions for TG and ABC
forcing. However, it is unclear whether this is related with
non-universal effects associated with interactions with the large scale
forcing, or with the use of the ESS hypothesis. For individual vortex
tubes in several flows, we verified that the flow inside and surrounding
the vortex tube is helical, as found before
\cite{Tsinober83,Moffatt85,Moffatt86,Levich87,Farge01}. Note that the
development of helical structures in a turbulent flow can lead to
the depletion on nonlinearity and a quenching of local interactions
\cite{Moffatt92,Tsinober}.

Concerning the energy transfer and triadic interactions, we confirmed
for several forcing functions that the cascade of energy is local between
Fourier shells, although it is strongly mediated by individual triadic
interactions which are nonlocal in nature. As a result, the energy cascades
from one shell to the next with a fixed step proportional to the forcing
wavenumber $k_F$. This effect was observed even in simulations using
random forcing with a correlation time one order of magnitude smaller
than the large scale turnover time. No qualitative differences have been
observed as the Reynolds number was changed.

However, the degree of nonlocality observed depends on the quantity
studied. A hierarchy is found in the relative amplitude of nonlocal
effects in triadic interactions, the shell-to-shell energy transfer,
and the energy flux. Triadic interactions are dominated by interactions
with the large scale flow. As more modes are summed to define the
shell-to-shell transfer and the flux, the larger population of small
scale modes starts dominating, and in the $1024^3$ simulations the
large scale flow is responsible for only $\sim 20\%$ of the total
flux.

An increase of the relative contribution of nonlocal interactions
(with the large scale flow and with modes outside the octave band studied)
in the total flux is observed as the dissipative range is reached. This
result is in good agreement with claims that the bottleneck effect is due
to the quenching of local interactions at the end of the inertial range
because of the presence of a cut-off in wavenumbers
\cite{Herring82,Lohse95,Martinez97}.

It is worth noting that the local energy cascade through nonlocal
interactions, or non negligible nonlocal interactions, has been observed
in the past in simulations at lower Reynolds numbers (see e.g. Refs.
\cite{Domaradzki90,Okhitani92,Yeung95,Zhou96}). Our results confirm
the presence of interactions between disparate scales in a turbulent flow
at much larger Reynolds numbers, and for a variety of forcing functions
and forcing scales. The results discussed here also shed some light on
the controversy in the literature about the relevance of the nonlocal
interactions. It has been claimed that interactions are local if
a different measure for the locality is introduced \cite{Zhou93,Zhou93b},
or if wavelets or a binning of Fourier space in octaves is used
\cite{Kishida99} (note that wavelets naturally introduce a binning in
octaves of the spectral space). The hierarchy found in the different
transfer functions is the reason for this apparent inconsistency
between previous results. As more modes are summed (e.g. to define
octaves in spectral space, or to define our partial fluxes $\Pi_P$),
the small scales overcome the triadic interactions with the large scale
flow, and local interactions give the largest contribution. This is also
in agreement with recent theoretical results about the locality of the
energy flux \cite{Eyink05}, or the locality of the shell-to-shell energy
transfer in Fourier octave bands \cite{Eyink94,Verma05}.

The fact that in simulations with $R_\lambda \sim 1000$ most of
the flux is due to local interactions, does not preclude however the
existence of strong nonlocal interactions with the large scale flow at
the triadic interaction level, and it should be kept in mind that even
at large values of $R_\lambda$ these interactions are responsible for
a non-negligible fraction of the total flux. The presence of nonlocal
interactions are a deviation from the standard hypothesis often
associated with Kolmogorov (K41) theory \cite{Kolmogorov41},
and can possibly explain departures from self-similar
models of the higher order structure functions and
controversial results observed in experiments, such as
the slower than predicted recovery of isotropy in the small scales
\cite{Shen00}.

Similar results were obtained for the helicity transfer. The
injection of net helicity in the large scales breaks down the mirror
symmetry in the flow. While confirming the direct cascade of helicity
\cite{Brissaud73,Andre77,Borue97,Chen03,Chen03b,Gomez04},
we also found a slower than
expected recovery of the symmetries in the small scales, with an
excess of relative helicity at small scales compared with the $k^{-1}$
expected drop. As in the energy cascade, the cascade of helicity takes
place in fixed steps proportional to the forcing scale, indicating
strong nonlocal triadic interactions.

Nonlocal interactions can also be responsible for observed
departures from universality. In \cite{Alexakis05b} it was shown with
a simple dimensional argument how nonlocal interactions can still
be consistent with a $k^{-5/3}$ energy spectrum. Vortex tube stretching
by the large scale flow plays a significant role, and the argument
has points in common with multifractal models of intermittency, such
as the $\beta$ model (see e.g. \cite{Frisch}). As previously mentioned,
in this work we showed evidence of a positive correlation of large scale
shear and small scale gradients, both in pdfs of velocity increments and
in structure functions. Ref. \cite{Laval01} showed that the anomalous
scaling of the structure functions is reduced when nonlocal interactions
with the large scale flow are artificially suppressed in a simulation.
Both results suggest that departures from K41 theory can be associated with
the imprint of the large scale forcing in turbulence.

The modern language used in turbulence is related to a great extent
to the K41 theory developed for the isotropic and homogeneous case, with
the assumption that such fundamental symmetries of the equations would
be recovered in the small scales even when broken (e.g. through external
forcing) in the large scales. To apply this theory to real turbulent
flows, the idea that small
scales restore isotropy and homogeneity is based strongly on the assumption
of local interactions between scales. The persistence of
anisotropies and other deviations from K41 observed in experiments and
simulations have been associated in this work with the presence of strong
non-local triadic interactions. At this point, the reader could ask
how much of the edifice of homogeneous and isotropic turbulence remains.

One of the assumptions in K41 theory is that the properties of the
inertial range are universal. This is directly related to the assumption
of local interactions. Since eddies in the inertial range only interact
with eddies of similar size, as the energy cascades through the inertial
range a self-similar solution is obtained. The fraction of the flux due to
non-local interactions, found to be $\sim 20\%$ here,  can
be interpreted as a subleading contribution to the local flux.
Deviations from K41 theory are well documented and several theories have been
proposed to explain them (see e.g.
\cite{Kolmogorov62,She94,Frisch,Lesieur,Tsinober}).
They have also been derived theoretically for the passive scalar
in the context of the so-called Kraichnan model \cite{Kraichnan68}.
Several of these corrections were associated with viscous effects for
turbulence at finite Reynolds number. What our results show is that
these deviations can come as well from interactions with the large
scale flow. In this context, simulations at higher resolution could help
to study the scaling of the non-local flux with the Reynolds number.

It is known that in the Kraichnan model only prefactors of the scaling
laws depend on the large-scale forcing (i.e. the exponents are universal:
independent of the forcing). However, in the case of the Navier-Stokes
equations, this has not been proved yet since the hierarchy of equations
for the $n$-point correlation functions is not closed. Then, inertial
range solutions with exponents that depend on the forcing
may exist.

In the case of realistic turbulent flows as encountered in astrophysics
and geophysics, the presence of strong non-local
interactions can lead to the persistence of anisotropies in the small
scales. In this case, the anisotropic effects question the applicability
of the theory of isotropic and homogeneous turbulence to real flows (at
least at Reynolds numbers comparable to the ones studied in this work).
The properties of the large scale flow can be more important than
expected to shape the small scales. A systematic study of the isotropic
and anisotropic contributions to the scaling
\cite{Biferale01a,Biferale01b,Biferale02} can be a first step to recognize
universal (and non-universal) features in these cases.

Finally, as previously indicated in \cite{Alexakis05b}, the existence of
non-negligible nonlocal interactions in a variety of turbulent flows
gives support to models involving as an essential agent of the nonlinear
energy transfer the distortion of turbulent eddies by a large-scale
flow (e.g. as in rapid distortion theory and its variants
\cite{Laval01,Dubrulle04}, or as in the Lagrangian averaged Navier-Stokes
equations \cite{Holm98,Chen98} where the turbulent flow interacts with a
smooth velocity field). We believe that the understanding of scale
interactions
in turbulence can lead to the development of a new generation of subgrid
models, beyond the usual hypothesis of locality done in most Large Eddy
Simulations.

\begin{acknowledgments}
The authors would like to express their gratitude to J.R. Herring for
valuable discussions and his careful reading of the manuscript. The
authors also acknowledge discussions with L. Biferale and S. Kurien.
Computer time was provided by NCAR and by the National Science Foundation
Terascale Computing System at the Pittsburgh Supercomputing Center. The
NSF grant CMG-0327888 at NCAR supported this work in part and is gratefully
acknowledged. Three-dimensional visualizations of the flows were done
using VAPoR, a software for interactive visualization and analysis of
terascale datasets \cite{vapor}. The authors are grateful to A. Norton
and J. Clyne (SCD/CISL) for help with the visualizations.
\end{acknowledgments}

redpajama_set_name: RedPajamaArXiv


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text:
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Health Promotion Major
What Does it Mean to Study Health Promotion?
The discipline of health promotion and education, also called wellness studies, aspires to educate communities in the skills needed to maintain personal health. As medical costs skyrocket, it has become clear that the health of individual Americans has an effect on society as a whole. The cost of treating preventable conditions is passed on to everyone via increased taxes and health insurance premiums. Health promotion and education professionals emphasize improvements in diet, exercise, behavior and even attitudes.
This process may seem simplistic, but students who take on health education as a career become familiar with the intricacies of physical health and the challenges of motivating a community or society toward awareness of health issues.
What Does a Health Promoter and Educator Do?
Health educators promote awareness of health issues as diverse as:
HIV/AIDS awareness and prevention
Dental disease prevention
Bone disease prevention
Teenage pregnancy prevention
Emergency medical practices
Health promoters and educators often work with the community. Health education and promotion programs generally focus on the health and well-being of individuals, instead of addressing theory or policy. The fields of study that serve as the foundation for health promotion and education include:
Inspiring Healthy Decisions
Motivating a society to actively encourage good health is no small task; Americans in particular are fond of junk food, sedentary jobs and low-key leisure activities. Improving public health takes legions of professional health educators working in every possible venue from schools to mass media. This career path includes such responsibilities as:
Organizing and motivating the community to work toward the common goal of overcoming health issues.
Utilizing technology and information to educate entire communities about promoting the health of each individual.
Organizing and implementing programs that promote healthy living.
Studying, maintaining and evaluating public programs that promote healthy living.
Organizing groups and coalitions that advocate health awareness in government and enact or change policies to this end.
Do You Have What It Takes to Be a Health Promoter?
Health education is a rigorous profession. Before embarking upon any course of study, make sure that this field is right for you. This field requires intensive study; it is a combination of science and social work.
Key Skills of Health Promotion Specialists
Professional health educators develop many of the following skills during their degree programs:
Assess individual and community needs. Health educators must be careful not to impose their own agenda on a community without first understanding the real wishes of its residents. Otherwise, their initiatives are likely to be met with significant resistance. Skilled educators can integrate surveys and polls into the first phase of a long-term program to test the waters for future outreach.
Plan and develop health education programs. Based on professional assessments, health educators work in teams to strategize long-term rollouts of their programs. By prioritizing program elements based on budgets and public demand, effective educators build awareness of their program long before they begin to supply key services in target areas.
Implement health education programs. After developing core strategies and building public demand, health educators must hit the streets to deliver their messages. Depending on the kind of program, educators may use a variety of mass media and direct client contact to execute their plans.
Manage programs and personnel. Skilled health education professionals oversee teams of specialists that work together to deliver program services directly to target communities. Team leaders often recruit publicists, community organizers, social workers, teachers, clergy, and other key people to execute successful programs. Juggling the schedules of these team members and supplying them with resources becomes the central focus of the project manager.
Evaluate health education programs. To justify continued funding and to gauge the effectiveness of programs, health educators develop a set of metrics they can use to quantify program results. Much like the process of evaluating customer service in a private business, program leaders must act quickly if programs fail to meet the objectives of the original plan.
Act to change government policy. Health education and promotion does not just take place in neighborhoods and in hospitals. Many health educators participate in grassroots lobbying efforts to inform lawmakers of the needs of their consitutents and the effects of health care legislation. Professional health educators meet with politicians, organize rallies, and inform voters about changes to health care policies.
Build coalitions. Even experienced health educators can have little impact without the support of community leaders and influencers. By involving community organizations, hospitals, politicians, research foundations, and other supporters, health educators can leverage the power of networks to accelerate the growth of their programs.
Identify resources and make referrals. Health education and promotion professionals are part of a network of people responsible for promoting wellness in their community. During the course of developing or administering programs, team leaders and staff members often encounter individuals who require more direct medical attention. Educators must teach their teams how to recognize cases of abuse, alcoholism or chronic illness, as well as what actions to take in each circumstance.
Develop audio, visual, and print materials. Though many health education programs focus on building strong presences in a community, health educators develop support materials to draw more people to live events and to support their overall learning objective. Some health educators rely on experienced production companies and ad agencies to develop multimedia campaigns. Other agencies and education specialists develop these materials in-house, using new technologies to save money and time.
Mobilize communities for action. The most successful health education programs empower community residents to take charge of their own lives long after the end of the campaign. Most professionals develop timelines for programs with a fixed end date to appeal to funding sources that might be reluctant to support open-ended efforts. Therefore, educators rely on community leaders to help take charge and keep the energy of the program building in local areas even after the professionals have moved on to new projects.
Train volunteers. Many health education programs, especially those that involve large, live events, rely on the "sweat equity" of community volunteers. These supporters come from all walks of life. Some are skilled professionals that lend targeted expertise to areas like construction or event production. Most volunteers simply want to help in any way possible. Program coordinators must anticipate training volunteers to learn essential skills to make the education program a success. Without effective training and supervision, volunteers can eat into time and resources meant to help community members.
Conduct research. Though many health education programs promote an existing agenda, they also present an opportunity to conduct important research into medical conditions in targeted communities. Health educators can often find additional funding for their programs by piggybacking their education efforts with research programs that can measure the current state of community health. As a result, researchers can return to the communities and more effectively measure the impact of the educational programs. Positive results often encourage educators and funders to replicate the programs in other target areas.
Write grant applications. Health educators rely on corporate backers, government agencies and private foundations to fund their programs. Because competition for funding can be fierce, many health education professionals learn how to write effective grant applications. It is important to project clear results that connect directly to the missions of the companies and organizations that funding is being requested from.
Write and present scholarly papers. Though health educators spend much of their time working with members of their target populations, experienced professionals understand the importance of reporting results back to the medical community. Therefore, usually as part of the final phase of their programs, team leaders coordinate the research and composition of journal articles and other scholarly papers. These reports remind doctors and other medical professionals of the importance of community outreach and education. They also create blueprints for other educators, helping them to replicate successful programs in their own neighborhoods.
Types of Health Promotion Degrees
Developing the strategies mentioned above can be fulfilling work. Professionals in this field are motivated by their desire to see a community take responsibility for its own well-being. When people understand the factors that affect their physical health, they become empowered to improve their habits and behaviors. In health promotion, knowledge is power.
College degrees in health education take many forms; some have clinical components, while others are more focused on marketing, community outreach and policy. Online degree programs in health education, particularly at the master's level, can be useful for working professionals who are ready to move to the next rung of the career ladder. Online schools offer a variety of concentrations in health education, from clinical programs in nurse education to the Doctor of Health Education degree program.
Associate Degrees in Health Education
An associate degree in health education might cover topics such as:
How to develop a fitness program
How to assess fitness
Human health and wellness
Associate degree programs generally require applicants to hold a high school diploma or GED equivalent, and to have earned at least a 2.0 GPA in lower-level math and science courses. AAS credits are often transferable to bachelor's degree programs, which may be required for career mobility after the student has gotten his or her foot in the door.
Bachelor's Degrees in Health Education
Coursework in a four-year bachelor's degree program is usually more in-depth and rigorous than that of an associate's degree program. The curriculum may include general education courses in the humanities and social sciences, as well as more specialized courses such as:
Skills in writing and oral presentations
Health and social theory courses
Developing leadership skills
Generally, earning a B.A. in Health Promotion does not require practical assessments and fieldwork. Institutions that offer degree programs in health promotion and education may also offer closely related degrees such as health studies, healthcare law and ethics, or health services management.
What Can You Do With a College Degree in Health Promotion?
The particulars of this career depend on your chosen specialty. Many positions require professionals to perform some or all of the following tasks:
Planning health education activities
Conducting workshops on health concerns
Identifying target populations
Developing public media campaigns
Conducting pre-program and post-program surveys
Performing screening services for individuals
Performing follow-up evaluations
Large organizations, such as the American Cancer Society or the American Diabetes Association, rely on health promotion and education graduates to pursue organizational missions. Non-profits generally employ educators to help design programs and strategies, give presentations on health awareness to communities, train volunteers, raise funds, and organize special events.
Clinical Settings
There is plenty of career crossover between health education and clinical practice, and many graduates choose to supplement their health education degree with clinical experience. This can be useful for community outreach programs that combine health education with basic services such as immunizations, checkups, blood tests and counseling.
Health educators enjoy a variety of opportunities in the corporate sector. Many corporations want to improve the health awareness of their employees. New research shows that health promotion programs enhance employee attitudes, reduce absenteeism, and lower stress. All these factors make for a better and healthier working environment while lowering healthcare costs.
Large, in-house programs typically employ health educators, while smaller companies are more likely to hire consultants for temporary workshops. Aside from outlining general practice guidelines that can assist the corporate community in improving its diet and exercise behavior, these programs often include individual needs assessments. This process determines what each worker can change to improve their stress management, weight management, recreation, or nutrition.
Emergency Medical Field
Health educators can also receive EMS training to become EMTs or paramedics. Some of the jobs available in this realm include work in hospitals, ambulance, fire service, law enforcement and clinical health care. Health education develops skills in public speaking, program planning, evaluation and implementation of strategies. The prevention of illnesses and promotion of health greatly affects emergency care professionals. Health education professionals are continuing to develop a stronger impact in this neighboring field.
Professionals wanting to work directly as educators may look to the public school system. The real work of health promotion begins with altering the perceptions and behavior of youth, for it is during the teen years that many lifelong health habits are formed.
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Q: component definition is missing display name on HOC I'm trying to create a higher order component but keep getting this eslint waring.

component definition is missing display name

I tried adding a display name like below but it still complains.
import React from 'react';

const HOC = props => (WC) => {
  WC.displayName = 'test'
  return (
    <WC />
  );
}

export default HOC;

A: Two things you need to correct.
First: Fix order of your functional component declaration.
Second: setting displayName to the component returned from HOC
const HOC = WC => {
  const MyComp = (props) => {
    return (
        <WC {...props} />
      );
  }
  MyComp.displayName = 'test'
  return MyComp;
}

Once you make the above change. You just need to use the HOC like
const MyCompWithHoc = HOC(CompA);

and render it like
<MyCompWithHoc propsA={'A'} {...otherPropsYouWantToPass} />

A: Your HOC should take the component as the first argument instead of props, see the example in the React docs: https://reactjs.org/docs/higher-order-components.html#convention-wrap-the-display-name-for-easy-debugging
EDIT: You can pass it props either by returning a functional component:
const HOC = WC => props => {
  WC.displayName = 'test'
  return (
    <WC {...props} />
  );
}

Or just return the wrapped component after setting the displayName:
const HOC = WC => {
  WC.displayName = 'test';
  return WC;
}

A: HOC is the component that's lacking a displayName property. The code you have above is changing the displayName property of WC which you don't want.
HOC.displayName = 'some higher component'

redpajama_set_name: RedPajamaStackExchange


text:
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Aleksey Zhukov
(1965-01-11) 11 January 1965 (age 54)
Volgograd, Russian SFSR
1.78 m (5 ft 10 in)
FC Rotor Volgograd 13 (1)
FC Torpedo Volzhskiy 28 (7)
FC SKA Rostov-on-Don 2 (0)
FC Spartak Ordzhonikidze 20 (0)
FC Torpedo Volzhskiy 152 (50)
Kemin Pallotoverit-85 ? (21)
RoPS 5 (1)
Kemin Pallotoverit-85 73 (9)
RoPS 27 (1)
FC Viikingit 111 (19)
SoVo ? (1)
FC Lootus Kohtla-Järve 40 (10)
FC Lootus Kohtla-Järve
Aleksey Zhukov (Russian: Алексей Жуков; born 11 January 1965) is a former Russian professional football coach and a player.
He played 2 games in the Soviet Top League in 1985 for FC SKA Rostov-on-Don. He then played in Finland for 16 seasons.
Player profile on Soccernet.ee
Aleksey Zhukov at FootballFacts.ru (in Russian)
This biographical article related to a Russian association football midfielder born in the 1960s is a stub. You can help Wikipedia by expanding it.
Retrieved from "https://en.wikipedia.org/w/index.php?title=Aleksey_Zhukov&oldid=931064367"
Sportspeople from Volgograd
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